Nuclear Physics 26 (1961) 5 2 9   5 7 8 ; ( ~ NorthHolland Publishing Co., Amsterdam Not to be reproduced by photoprint or microfilm without written permission from the publisher
T H E O R Y OF N U C L E A R R E A C T I O N S I. R e s o n a n t S t a t e s a n d C o l l i s i o n M a t r i x J. H U M B L E T *
Institute/or Theoretical Physics o/ the University, Copenhagen and Cali/ornia Institute o/ Technology, Pasadena and L. R O S E N F E L D B I O R D I T A  Nordisk Institut /or Teoretish Atom[ysih, Copenhagen l~eceived 15 M a y 1961 A n e w f o r m u l a t i o n of t h e g e n e r a l t h e o r y of n u c l e a r r e a c t i o n s is proposed, o n t h e b a s i s of a definition of t h e r e s o n a n t s t a t e s as d e c a y i n g s t a t e s c o r r e s p o n d i n g to c o m p l e x e i g e n v a l u e s of t h e t o t a l e n e r g y of t h e c o m p o u n d s y s t e m . A c h a r a c t e r i z a t i o n of t h e s e s t a t e s is d e r i v e d f r o m " n a t u r a l " b o u n d a r y c o n d i t i o n s e x p r e s s i n g t h e a b s e n c e of i n c o m i n g w a v e s in all c h a n n e l s : s u c h b o u n d a r y c o n d i t i o n s yield in a unified f o r m t h e b o u n d s t a t e s , t h e p r o p e r r e s o n a n t s t a t e s a n d " v i r t u a l " s t a t e s of a t y p e familiar f r o m t h e case of t h e 1S s t a t e of t h e d e u t e r o n . T h e dispersion f o r m u l a e for e l e m e n t s of t h e collision m a t r i x d e t e r m i n i n g t h e r e a c t i o n crosssections a r e established, as i m m e d i a t e c o n s e q u e n c e s of t h e i r a n a l y t i c a l b e h a v i o u r as f u n c t i o n s of t h e c o m p l e x e n e r g y variable, b y a p p l i c a t i o n of t h e MittagLeffler t h e o r e m o n t h e rep r e s e n t a t i o n of m e r o m o r p h i c f u n c t i o n s . T h e p r e s e n t a t i o n t h u s o b t a i n e d , besides b e i n g s i m p l e r t h a n t h e u s u a l ones, h a s o v e r t h e l a t t e r t h e a d v a n t a g e of e l i m i n a t i n g e v e r y a r b i t r a r y e l e m e n t f r o m t h e specification of t h e r e s o n a n c e s a n d of t h e n o n  r e s o n a n t b a c k g r o u n d ; t h e s e e l e m e n t s of t h e d e s c r i p t i o n are, in p a r t i c u l a r , i n d e p e n d e n t of t h e choice of t h e c h a n n e l radii, a n d a p p e a r a l t o g e t h e r as i n t r i n s i c p r o p e r t i e s of t h e c o m p o u n d s y s t e m .
Abstract:
Introduction Since the first theories of nuclear reactions were proposed some twentyfive years ago, numerous papers have been written with the purpose of justifying the existence of formulae of the dispersion type for the cross sections, which have been found well suited to represent the experimental data. In view of the obvious insufficiency of perturbation treatments 1), special methods had to be devised, involving in particular a definition of the resonance levels. In spite of the progress made in this direction b y successive theories, none of these can be said to be satisfactory from the physical point of view, inasmuch as the definitions of the resonances they propose contain arbitrary elements. The Rmatrix theory developed b y Wigner and his collaborators ** introduces the resonances in a very indirect and complicated way. It is particularly regrettable that a central role be assigned to such a quantity as the Rmatrix, which has no direct physical meaning; when simplifying assumptions or * O n l e a v e of a b s e n c e f r o m t h e U n i v e r s i t y of Liege, B e l g i u m . t t F o r references, see t h e review article b y L a n e a n d T h o m a s 3). 529 September 1961
~30
J. H U M B L E T AND L. R O S E N F E L D
approximations are introduced, they must consequently concern the Rmatrix itself, and m a y still leave the collision matrix in a very complicated form. One should also remember that in the Rmatrix theory one has to introduce a number of quantities which are essentially dependent on the "channel radii". This is in fact the case with the Rmatrix itself, whereas a physical quantity such as the collision matrix does not, of course, show such a dependence. The definition of the resonances which had been proposed earlier b y Kapur and Peierls 3) avoids most of the complications of the Rmatrix formalism, but is nevertheless very artificial, since the positions of the resonance levels it introduces are explicitly dependent on the energy of the incoming particles. It is our purpose, in several papers of which this is the first, to develop a theory of nuclear reactions free from such arbitrary features, and ill which the resonances directly appear as intrinsic properties of the compound system. The possibility of setting up a "natural" definition of the resonance levels was indicated long ago b y Siegert 4) and used in an earlier paper 5) to establish a rigorous dispersion formula for the special case of elastic scattering b y a fixed potential field of finite range *. This dispersion formula was derived from a MittagLeffler type of expansion of the scattering amplitudes. Recently, Peierls ~) and Le Couteur s), using the same natural definition of the resonances, have made interesting contributions to the analysis of the structure of the dispersion formulae when several channels are involved. However, Peierls uses the less general Cauchy type of expansion, while Le Couteur is mainly concerned with the expansion of the collision matrix as an infinite product. On the other hand, a direct extension of the method developed in the case of elastic scattering ~) to more general types of nuclear reactions proves possible without essential complications. An outline of the main physical ideas enabling such an extension under neglect of spins and Coulomb effects has been given b y one of us in several courses of lectures during the last few years: the most recent of these 9) forms the starting point for the more complete theory developed here ft. Before treating the general case of nuclear reactions in the second part of this paper, we have found it convenient to devote a first part to a renewed discussion of the elastic scattering. In sections 1 and 2, the main results obtained earlier 5) for the elastic scattering of neutral spinless particles are briefly summarized and some new results added; the accent here is put on the physical interpretation of the methods employed rather than on their mathematical justification, which is fully discussed in ref. 5). Section 3 deals with the elastic t F o r this case, one should also consult a review article b y N e w t o n e), where f u r t h e r references will be found. I n a f u r t h e r p a p e r (preprint), N e w t o n extends some of his results to more general problems. t? An a b s t r a c t of the p r e s e n t p a p e r h a s been s u b m i t t e d to the K i n g s t o n Conference on Nuclear S t r u c t u r e 10).
THEORY
OF
NUCLEAR REACTIONS (I}
531
scattering of charged spinless particles, which was not discussed in ref. 5). The first part must be regarded as an introduction to the notations and methods used in the general case o~ nuclear reactions. Many of the properties of the wave function are more easily recognized in the case of elastic scattering, and later we shall only have to justify their extension to the general case. In this paper, the theory of nuclear reactions will be limited to the analysis of the properties of the collision matrix and to its expansion. The corresponding properties and expansion of the cross sections will be discussed in another paper. However, in the case of elastic scattering the cross sections have also been written down in order to facilitate the physical interpretation of the parameters introduced in the expansion of the scattering matrix.
P a r t I: E L A S T I C S C A T T E R I N G OF S P I N L E S S P A R T I C L E S 1. N e u t r a l P a r t i c l e s : S c a t t e r i n g M a t r i x and R e s o n a n t S t a t e s 1.1. D E F I N I T I O N
OF THE
SCATTERING
MATRIX
Let us consider the elastic scattering of neutral spinless particles b y a target nucleus and assume that the interaction m a y be represented b y a real potential V which is spherically symmetric in the barycentric system of reference. If the method of partial waves is used, the solutions of the Schr6dinger equation H ~ = E ~ , for a total energy E, take the form, in polar coordinates r, f2 ( ~ 6, ~, the polar axis being the direction of the incoming particles),
~=
X~
_
1 V~
~
~ / 2 ~  ~ i ' P , ( c o s O ) uz(r~)
~=o
(1.1)
r
The radial wavefunction u, is a solution of the equation d~
Ur#+ k~
l(l+l) r2
)
v(r) uz(r, k) = O,
(1.2a)
satisfying the condition uz(0, k) = 0.
(1.2b)
In these equations we have introduced the notations
k 2 = 2ME/ti e
(1.3)
v(r) ~ 2MV(r)/]i 2,
(1.4}
and where M is the reduced mass. It will be convenient to call k 2 an energy and v(r) a potential. We assume that the potential v(r) is everywhere finite, and that it vanishes identically at all distances larger than some finite distance a:
v(r) ~ 0 for r > a;
(1.5)
~32
J. HUMBLET AND L. ROSENFELD
the limiting distance a is normally of the order of magnitude of a nuclear radius, although it can be taken much larger if necessary, as we sh~dl see later. It appears that the restriction 0.5) on the potential   excluding an exponentially decreasing behaviour, for instance   is quite essential for the mathematical simplicity and generality of the following discussion. On the other hand, it clearly does not imply any serious physical limitation (cf. section
3). For r > a, solutions of eq. (1.2a) are the wellknown spherical Bessel and Hankel functions defined b y
y , ( x ) = v'½~x],+~(~), ~,(1)(X)
=
V17gX /  / ~ t ( ~ ) ,
Y_,_l(x) = V1~xJ_,_~(~),
(1.6)
fft(2'(~)  %/Izcx Hl~½(x),
(1.7)
where x = kr. They satisfy the relations J/d~(1) +~f,(~) = 2 J ~ ,
~e,( = ] ,   i (   1 ) ' J _ H ,
(1.8)
~,(~) = ] , + / (   1 ) ' ] _ , _ ~ .
(1.9)
It is useful to recall the asymptotic expressions 2'l! o~f (1) ~ i(t+l) ei~,
5~, (x~ ~   i (2l)! x' (for small x),
(1.10)
~fl(2) ~ i~+1 ei~
(1.11)
(for large x).
For r > a, we choose two linearly independent solutions of eq. (1.2a) O,(r, k) = i ~ z ( l ~ ( k r ) ,
It(r, k) = i~Y,(2~(kr),
(1.12)
whose asymptotic behaviour for kr >> 1 is O~ ,~ eqk~½~),
I~ ~~ e~(k'½~'),
(1.13)
and which therefore respectively represent an outgoing and an incoming spherical wave. Consequently, for r > a, the solution of eq. (1.2a) m a y be written as a superposition of outgoing and incoming waves of appropriate amplitudes xz, y, in the form (1.14)
u~ = x~O,+y~I~.
For reasons that will appear later, O, and I, have not been normalized to unit flux. The exterior solution (1.14) must be joined smoothly to the interior solution nz, which embodies the effect of the potential field. Using the notation
1'(,, k) = 1'(,) 
T, 1(,, k),
1'(,,, k) = 1'(,,) 
1(,, k) ,=,
Ta~o~Y oF NUCLEAR=ACTIOSS (I)
53a
we write down the conditions expressing the c o n t i n u i t y of u s and u'~ for r = a:
us(a, k)  xzOt(a, k)+y~It(a, k),
(1.15a)
u',(a, k) = x~O'z(a, k)+yzI't(a, k).
(1.15b)
F r o m these we m a y now derive an expression for the amplitudes x~, yz in terms of the interior radial wavefunction u t and its derivative. It is convenient to m a k e a systematic use of the notion of Wronskian of two functions / and g,
W(/, g; r, k) /(r, k)g'(r, k)/'(r, k)g(r, k);
(1.16)
this Wronskian will also be written more briefly as W(/, g) or W(], g; k) when no confusion is possible. It is well known t h a t the Wronskian of ~,%f(1) and 3~, (~) is independent of r: W ( W t (1), ~t(~); r, k) = 2ik,
(1.17)
w(o , _rt; r, k) =  2 i k .
(1.1s)
or, b y eqs. (1.12),
We thus get from eqs. (1.15) 1
x~ 
2ik W(u , It; a, k),
Yt = ~
1
W(ut, 0~; a, k).
(1.19a) (l.19b)
The ratios
xt W(uz, It; a, k) Yt W(uz, 0 5 a, k)'
V~ =
(1.20)
for the different values of l, are the elements of the scattering matrix (which is diagonal in the v e r y simple case considered here). The scattering cross sections obviously depend on these quantities Uz a n d the values i V~~+
yz(un) :
1)
(1.21)
of the amplitudes corresponding to an incoming flux of one particle. In fact, taking account of eq. (1.21), the differential and integrated cross sections are readily expressed in terms of U t in the form de(Q) = ~
1
I ~ ( 2 / + 1)Pt (cos 0) (1gt)[2dQ,
(1.22)
7~
a  2~ a r t
2~ (21+i)[1V~l 2.
(1.23)
534
j. aUMBLET AND L. ROSENF~LD
1.2. P R O P E R T I E S
OF T H E SCATTERING M A T R I X
Our task is now to study the behaviour of the scattering matrix element U~ as a function of k, especially with a view to relating it to the observed resonance features of the cross sections. In this investigation we shall clearly have to treat k as a complex variable. Let us first determine somewhat more precisely the properties of the radial wavefunction uz (r, k) which enters into the definition (1.20) of Uz. Since the physical quantities (1.22), (1.23) only depend on the radial function uz through the matrix element U,, we see that the normalization of u, need not actually be related, as it would seem from eq. (1.19b), to the particular value (1.21) of yz which occurs in the expression for the cross sections: this normalization m a y be fixed arbitrarily. For our case of a potential everywhere finite, a convenient choice, satisfying the requirement (1.2b), is that defined b y the condition (1.24)
u,(r, k) _ 2'l! limr_,0 r ~+1 (2l+ 1) .v"
It is well known 5, s) that with this normalization u~ and u'~ are integral functions of k ~ for any positive and finite value of r. Further, the potential v(r) being real, uz and u' z are also real functions of k 2, so that u~(r,k) = u,(r, k),
(1.25a)
u~(r, k)* = u~(r, k*) = uz(r,k*);
(1.23b)
similar relations being also valid for u'z. From the properties $ff c1~(kr) ~ ( 1)z+lJ~z~2~ (kr), #ff ~l~(kr)* = ~
~2~(k'r)
(1.26a)
(1.26b)
and eqs. (1.12), one gets
O,(k)
= (1)ZL(k),
Ot(k)* = It(k* ) ~ (1)z0t(k*).
(1.27)
The definition (1.20) of Ut then immediately leads to the relations U#)U~(k)
= 1,
v,(k)* = v ~ (  k * ) ,
(1.28) (1.29)
whence =
1.
(1.30)
For real k, eq. (1.30) expresses the unitarity of the scattering matrix: Ul(k)*U~(k ) = 1
(for real k),
(1.31)
from which one derives for the partial cross section az defined b y eq. (1.23)
(I)
535
2z at = ~ (2/+1) Re [1Ut(k)].
(1.32)
THEORY" O F N U C L E A R
REACTIONS
the equivalent, and somewhat simpler, expression
1.3. DEFINITION
AND
FORMAL
PROPERTIES
OF THE
RESONANT
STATES
The characteristic parameters determining the resonance levels, their positions and widths, must be mathematically defined by some additional boundary condition imposed on the quantity us(a, k) on which the matrix element U~ depends. Considering that the resonances correspond to maxima of the cross sections, the expression (1.20) for U t suggests as a suitable resonance condition the equation W(u,, 0t; a, k) = 0. (1.33)
Regarded as an equation in k, it indeed defines, for each value of l, a set of (complex) roots kzn, some of which we m a y expect to be associated with resonances. Eq. (1.33) is the "natural" boundary condition first introduced b y Siegert 4). Let us start b y discussing its consequences from a purely mathematical point of view, and then show that it indeed provides an adequate representation of the physical situation. The roots ktn of eq. (1.33) m a y be described as complex eigenvalues of the radial wave equation (1.2), with corresponding eigenfunctions u,n(,)
(1.34)

For each kin we have identically W ( u z,
0 5 a, kt~ ) = 0.
(1.35)
The complex quantities kt~ will be written in the form kz~ = Kt~i~z~,
(1.36)
where Kt, and ~'t~ are real and where the minus sign has been chosen because, as we shall see presently, ~t~ is never negative when Kz. :/: 0. We shall often omit the index l in kt., Kz. and ~ . . Let us now discuss the location of the eigenvalues k. in the complex kplane. On account of eqs. (1.25b), (1.27), one has, on taking the complex conjugate of eq. (1.35), W(ut, Or; a,k*,,)  0; (1.37) this shows that if k~n = Kz,,i~tn is a solution of the boundary condition (1.33), the same is true of k*,, =
  J¢,,,iVt,,.
(1.38)
Without any restriction, we m a y therefore always assume that
K~. __> 0.
(1.39)
5~6
j . HUMBLET AND L. ROSENFELD
Starting from the wave equation (1.2) and the boundary identity together with their complex conjugates, we can easily prove that
]u,,,(a)]*[L,(kn)L,(kn)* ]
(k'*2k~) fo lu*~(r)l~dr'
(1.35), (1.40)
where L,(k) 
~/~"(ka)
~aZ{1) ( k a )
'
L~(k)* = L~(k*),
(1.41)
the last equation being a consequence of eqs. (1.26). Using the recurrence relations between the spherical Hankel functions, we can show 5) that
fw/1, (kn a)12EL, (kn) L, (k~)*l = i(kn + kn*)IWo '1' (kn a)12+ (k2k~ *~) W,~ (a), (1.42) where the quantity ~z~(a) is zero for l = 0 and, for l ¢ 0, is given by the essentially real and positive expression ~,,
Ca) = lak~l~{(2l 1) IX~I~)~(k,,a)7 _~_ (2l_5)[j,~l_)a(kna)12 + . . . _~_(2± 1)[3~+½(k~a)[~}, (1)
{1.43)
the 4 sign corresponding to the parity of l. For k,,+k~* = 2K~ ¢ 0, eqs. (1.40) and (1.42) yield
lu,,,(a)/O,(a, kn)lel e'~ °1 ~
=fo"tu*~(r)12dr+ )O,(a, u,.(a)kn)
(1.44)
showing that 7~ is positive when K,, is not zero. Consequently, the eigenvalues k, = K,i~,n for which y, < 0 (upper kplane) necessarily have K~  0: they lie on the imaginary axis, and they clearly correspond to the real (bound) levels of the system, since their energy   y ~ is negative and their radial wave function has the asymptotic form exp ( i k j ) exp ([wnlr). In the lower kplane the eigenvalues are symmetrical with respect to the imaginary axis or lie on it. It was shown in ref. 5) that there is no solution of the boundary condition (1.33) on the real axis, except possibly at k = 0, but we shall disregard this very special case ill what follows. In ref. 5) it was also proved that the number of solutions obeying the boundary condition (1.33) is infinite and that only a finite number of them can be of the type k. iy~ (7~ positive or negative). Another important property of the eigenvalues ks, is that they are independent o/the limiting distance a. This must be understood in the following way: 1) if v(r) = 0 for r > a, then, for b > a, the two boundary conditions
W(u,, 0,; a, k) = 0
(1.33)
T H E O R Y O F N U C L E A R R E A C T I O N S (II
537
and
W(u,, 0z; b, k) = 0
(1.45)
have exactly the same set of roots ksn; 2) if Iv (r) ] is sufficiently small between a and b, the roots of (1.33) and (1.45) whose moduli are not too large are nearly equal; this m a y even be true for b ~ oo under suitable conditions s). The eigenvalues introduced in the Rmatrix formalism 2) and ill the KapurPeierls theory 3) have no such property. Here it is simply connected with the fact that the boundary condition (1.33) has the form of a Wronskian put equal to zero. From the equation (1.2a) and that satisfied b y Os(r, k), (d 2
+ k2
l(l+ 1)~ ~ ] O~(r,k) = 0,
(1.46)
we easily derive, b y a wellknown procedure, the relation
W(u z, 0,; b, k) = W(u v 0l; a, k) f~ v(r)ul(r , k)Os(, , k)dr,
(1.47)
which proves the announced property of the eigenvalues k,,. From eq. (1:47) and the similar one with I z instead of 0 l we see that the scattering matrix UI, which is the ratio of the two Wronskians, is itself "independent of a". 1.4. EXPANSION OF THE SCATTERING MATRIX
We noticed earlier that the functions u s and u' s are integral functions of kS; obviously, k ~01, k~Iz and their derivatives are likewise integral functions of k. Consequently U s is a meromorphic function of k. We shall assume here that its poles kn are all simple. We m a y then expand U s b y means of the following theorem, which is a straightforward consequence of the MittagLeffler theorem: Let ](k) be a meromorphic function of k whose poles k~ are all simple and such that 0 < Ikll =< Ik2[ _<_ . . . . and let p~ be the residue of ](k) at k : such that the series
0.48)
k~ ; if there exists an integer M ~ 0
Ip~l
(1.49)
]knlM+I is convergent, we have /(k)
=
_
kko'
(150)
where e(k) is an integral function of k, while the series is absolutely and uniformly convergent for all values of k uniformly distant from the poles kn.
538
J . H U M B L E T AND L. R O S E N F E L D
In applying this theorem to the scattering matrix element Uz, the most delicate point is the proof of the convergence of the series (1.49)and the determination of the least value of M ensuring this convergence; in fact, this requires a study of the asymptotic distribution of the eigenvalues k~n in the complex kplane. For a fixed potential field of finite range, this investigation has been carried out in ref. s). it is found that one must take M _>_ 1. To justify completely the application of the theorem (1.50) to the expansion of U~, we still have to verify, according to the first inequality (1.48), that k = 0 is not a pole of Uz. Indeed, from the definition (1.20) of U,, with eqs. (1.12) and the relation (1.8) between Bessel and Hankel functions, we have
W(u~, J,; a, k) U,1  2 W(uz,5~fz(l~; a, k)' eqs. (1.10) then yield, for small
ka,
(1.51)
the asymptotic expression
 2 i ( 2,z! ~* (ka),,÷l au',(,, 0)(Z+ 1)u,(~, 0) \(2~.! au'z(a, O)+lu,(a, O)
Uz 1 ,~ 2/+~
(1.52)
As we have excluded the exceptional case of a resonant level in k = 0, this formula shows that k = 0 is actually a zero of order 2 / + 1 of U~1. The residue p,(k,,) of U, is easily evaluated (its explicit expression m a y be found in ref. 5)); for p~(k*,) one has s)
p~(k*~) = p,(k,~)*.
0.53)
Ordering the eigenvalues kz, (and k*,) according to increasing moduli, one m a y therefore write for Uz1 expansions of the general type *
( UR,. (~)~ ~'MR*"~ V,(k)1 = " % ( k ) + k ~' .:i z_ \k~, k+k~,] where MQ, (k) is an integral function of k, while the coefficient MR,n is
(1.54) defined
by
~R,n
=
,,.p,(k,.)/k,L
,1 ezn
: 1½
for
~,. ~ 0 ,
for
Ktn ~ O.
0.55)
Since the property of U~(k) expressed b y eq. (1.29) is exhibited b y the term involving the series in the expansion (1.54), it must also belong to the integral function MQz(k):
mQ,(k)* = MQt(k*).
(1.56)
• I n t h e c o m p a r i s o n w i t h ref. 5) one s h o u l d be a t t e n t i v e to s o m e differences in t h e n o t a t i o n s ; t h e m a i n e x p a n s i o n s in ref. 5) were w r i t t e n for t h e a m p l i t u d e of t h e s c a t t e r e d w a v e ~ ( U t   1 ) .
THEORY
OF
NUCLEAR
REACTIONS
(i)
539
As none of the poles kz,,  k*, lie on the real kaxis, the expansions (1.54) are valid for all real values of k, which are those of physical interest. For such real k, the properties (1.29), (1.56) imply t h a t the real parts of Ut and MQ~, as well as t h a t of the series term, are even functions of k, i.e. functions of the energy E. According to eq. (1.32), this is therefore also the case for the cross sections a t. Since, according to eq. (1.52), MQt(O) = 0, we m a y also write for real k and E 2 Re MQz(k ) : k 2 MC~(E), (1.57) where MC~(E) is a function of E which is finite for E : 0. In particular, let us consider the case when the smallest value of M is adopted. Introducing the notations Q5 = 1Q~,
~, = l e t ,
~tn = 1Rln,
(1.58)
we have the following expansion of the cross section:
x R,_. 1
a t = u ( 2 1 + l ) I C , ( E )   4 Re , k 2 k ~ , , j .
(1.59)
The last t e r m m a y be p u t into a more explicit form b y means of the following notations:
(1.6o)
1 " ~2 k 2 :~ 8tn = Ez,,~*Fz.,
2M t,, i.e. ~2
Et.
~
~2 2 ~ (Kz.ytn),
F~,,

2M
4K~n~t n
(1.61)
and 2~ 2 
2~ 2
~  R t . = e,. ~
Pt. ~  x l t . + i B , . .
(1.62)
We t h e n have
~,(E) = =(2~+1) [C,(E) +.2"1 ~ {r,.B,.+(EE,.)At.I~~ J"
(1.63)
Another form of expansion for the cross sections az is of course given b y the formula (1.23), which, together with eqs. (1.54), becomes
0,(k)+ . = , (\kk,n R,.
R*. ], ' + k+k*~,,]]
(1.64)
where 1
~t(k) = ~ 1Qt(k)
(1.65)
is an integral function of k. Here again, eq. (1.29) m a y be used to show t h a t
J. HUMBLET AND L. ROSENFELD
540
the last factor of eq. (1.64) is a function of k ~ or E, but this expansion is not well adapted to bring out the Edependence explicitly. It is also important to observe t h a t none of the above expansions is appropriate to exhibit the behaviour of at at small energies. According to eq. (1.52), the latter behaviour is obtained by taking M = 2 l + 1. Both parts of the expansion (1.54) then have, like Uz1, a zero of order 2l+1 at k  0. With the notations
Qz(k) :
Rzn =
k2~+1 ,
2Z+lR~n _
k,~,
(1.66)
we obtain, instead of eq. (1.64),
+ ~ \kk~ + k+k*j
(1.67)
It is also easy to derive an expansion of the type (1.63) showing the smallenergy behaviour explicitly; it suffices to take M = 4 / + 1 ; then, although 4~+lQz has only a zero of order 2 / + 1 in k = 0, the real part of this q u a n t i t y is, for real k, a function of E which vanishes like E ~+1 at E = 0. The two terms of the expansion of ~ so obtained are then proportional to E 2~. Finally, it is useful to observe here t h a t the two special cases of the expansion (1.54) which have been used to derive eqs. (1.64) and (1.67) m a y be obtained directly b y application of the theorem (1.50) with M = 0 to the quantities
1 S~  ~ (U~I)
and
1 S t  k ~+1 (Usl),
(1.68)
respectively.
2. Neutral Particles: Physical Interpretation 2.1. P H Y S I C A L P R O P E R T I E S
OF THE RESONANT STATES
We shall now use the formulae just established for the cross sections to discuss the physical meaning of the complex eigenvalues k~n introduced by the condition (1.33), and of the real quantities Ez~ and F~n derived from them b y means of eqs. (1.60), (1.61). The very form of the donominators in the sum over n in the cross section (1.63) shows that E ~ and T'~ (if t h e y are positive) must be associated with the position and width of a resonance level of the compound system formed by the incoming particle and the target nucleus. Although the concept of resonance has its full meaning only when nuclear reactions are treated as manybody processes, we shall nevertheless introduce it here because the main features of resonance are already present in the elastic scattering, as far as the mathematical aspect of the problem is concerned.
THEORY OF N U C L E A R REACTIONS (I)
541
This will in fact lead us to a perfectly natural physical formulation of the mathematical problem involved in the theory of nuclear reactions. Moreover, the following considerations m a y describe an actual physical situation if it is assumed that the potential v (r) is strongly repulsive for r just smaller than a. According to eq. (1.19b), the boundary condition (1.33) corresponds to a physical situation in which there is no incoming wave. We know that this is the case for a bound state (since for such a state an incoming wave, proportional to exp (ikj) = exp (17.(r), would increase exponentially with r), and it is quite natural to take this property of "no incoming wave" as characteristic "also for the resonant states associated with a compou_ud system of positive energy. This point of view is exactly t h a t adopted by Gamow in his theory of adecay. Of course, the resonant states defined in this way, in contrast to the actual bound states, will not be completely stationary; t h e y must decay in time according to an exponential law.
2.1.1. Resonant states Let us first consider the physical situation corresponding to a proper resonant state kz~, i.e., one for which Kt~ > ~z~ > 0; both Ez, and Fzn are then positive. In this state, the asymptotic form of the wave function for large r, including the time factor, is proportional to exp (ikz~ri#z~t/~), i.e. exp [i(K,,
/',.t~
rE,,~t/~)~exp ( 27~ / exp (rz.r),
(2.1)
where Ezn and Ftn are defined by eqs. (1.61). The first factor of the product (2.1) represents an outgoing wave of wave number Kz, associated with a state of positive energy E~=. The second factor shows t h a t this state decays at the rate Ft,~/~, which is in conformity with the usual interpretation of the width ,of a resonance level. Since the velocity of the outgoing particle is
V~n J~Kt" M'
(2.2)
we see that the exponential increase with r of the last factor of the product (2.1) m a y be interpreted, in a familiar way, as due to the fact that, at a distance r, one finds those particles which left the decaying centre at a time trive, when the amplitude there was larger by the factor exp \ 2~
= exp (yz, r).
We must now discuss the physical situation corresponding to the solution k*, = Kz,iTz . of the boundary condition (1.33). By changing, in the
5~2
j.
H U M B L E T A N D L. R O S ~ N F E L D
product (2.1), KZn into Kzn, and consequently Fz~ into Fzn, we obtain exp [i(K~r + Ez~t/t~)lex p (Fz~t~ \ 2/~ ] exp (Tznr),
(2.3)
where E ~ and F ~ are still positive. We now have an incoming wave whose amplitude increases with time. This result means that the eigenfunction corresponding to k*n is simply related to that corresponding to k~ by time reversal. Indeed, a purely decaying state is characterized just as well by putting equal to zero the amplitude of the incoming part of the wave function 7 t, or the amplitude of the outgoing part of the timereverse ~b = K ~ . Now, if the Hamiltonian operator is invariant for timereversal, and the energy E complex, the function q~ is a solution of the equation H # = E * # . The resonant state is therefore defined by the condition W(uz, It; a, k%) = 0
(2.4)
as well as by
W(uz, 05 a, kin ) = O. We see from eqs. (1.25), (1.27) that these two conditions are identical. The former will give a purely ingoing #z = KSVz wave and the latter a purely outgoing hut wave. Consequently, we must consider kz, and k*~ associated not with two different resonant states, but with a single one only. This conclusion explains why in eq. (1.63) we could associate k~ with k*~ in a single resonance term. 2.1.2. Bound states and virtual states Let us now discuss the levels k~ for which E ~ = 0, i.e., [~z~[ ~ K~ >= 0. It m a y be t h a t there is no level of this type associated with a given potential v(r), but examples of the actual occurrence of such eigenvalues are known. Anyway, their number is always finite because, for large Ikz~l, i.e. large n, we have 5) Kzn : O(n) and ~ = O (log n). Firstly, we m a y have eigenvalues corresponding to bound states. Then   ~ , = Im kz~ > 0, Kz~ = 0 and therefore E~ = ~..
Fz. = O.
For l = 0, for instance, the corresponding term in the cross section (1.63) is ~A0,
~Ao.
EEo~  E + [Eo,,[ ;
(2.5)
it takes its maximum value for E = O. It is well known that the aS bound state of the deuteron gives a contribution of this type to the protonneutron scattering cross section at low energies.
THeOrY OF ~OCLEAR~EACT~O~S(~
54S
Secondly, we m a y have negative Ez~ with positive 7~, if {~ = arg k~ _<_ ~ , so that 0 <_ Kn =< ~ and hence Ez, _<0,
F~n > 0;
arg (k~*) is then in the interval (~~, ~~) . An example of such a level is also provided b y the protonneutron system, whose ~S state is of the type k~ = i7~(7,~ > 0); its contribution to the cross section is again an expression of the form (2.5). More examples of levels which are not bound although E ~ < 0, and additional information about their physical meaning, will be found in the recent papers by Nussenzveig n) and by Beck and Nussenzveig ~). To distinguish them from the proper resonant states, such levels m a y be called virtual; the corresponding terms in eq. (1.63) have to be added to those associated with the true resonances in order to give a complete expression for the cross section. Of course we could also suppose these terms to be removed from the resonant part and added to the nonresonant part C~. However, we shall not do so, at least when exact formulae are concerned, because this would destroy the analytic definition of this separation into resonant and nonresonant part. Moreover, there is experimental evidence 1~) in favour of negative energy "resonances" of the type corresponding to virtual states with F ~ > 0; t h e y give rise to a very pronounced increase of the cross section when the energy E goes to zero. 2.1.3. Potential scattering According to its mathematical definition as integral function, the nonresonant term Qz is a smoothly varying function of k. We m a y consider it as corresponding to the wellknown potential scattering term present in other theories, although it is not here related to tile scattering by a hard sphere of radius a. On the contrary, it has the same independence o[ a as the positions o] the levels k~ themselves. This is not the case when, instead of expanding according to the MittagLeffler theorem, one applies the Cauchy theorem 1~) to expand the product U 0 e 2~k~, or more generally UzO~(a, k)/Iz (a, k). For this reason, we have abandoned the use of such expansions, which were considered in refs. ~, 15). Because one m a y choose different values for the exponent M, the separation between potential scattering and resonance scattering is not unique; each choice of M corresponds to a different separation of the two contributions. In practice, this ambiguity (which is obviously inherent in the whole conception of the resonance process and quite general) might be used to facilitate the fitting of the theoretical cross section to the experimental data: it gives more freedom than the usual definition of the potential scattering to adjust the "background"
544
J. HUMBLETAND L. ROSENFELD
against which the resonances are determined. If one is concerned with the b e h a v i o u r of the cross section at small energies, there is, as we have seen, a definite choice of M which makes this b e h a v i o u r explicit. 2.1.4. Case o] no potential I n concluding this discussion, we w a n t to point out t h a t a v e r y strong a r g u m e n t in f a v o u r of the definition of the resonant states a d o p t e d here is t h a t it gives no resonant states when there is no potential. This is a consequence of the fact t h a t the corresponding function u z is simply J~(kr)/k z+l, for which 1
W(u~, 0~; a, k) 
k~.
I n o t h e r formalisms 2, a) the b o u n d a r y conditions a d o p t e d define resonant states also for v (r) ~ 0, which is of course less satisfactory. 2.2 THE BREITWIGNER APPROXIMATION It is easy to p r o v e 6) t h a t if a resonant state k. is such t h a t 7, << IK,K,~I[ a n d if 2Kn is of the order of
Kn+l~Kn_l, it
(2.6a)
m a y be considered " n a r r o w " , i.e.
7~ << K~,
(2.6b)
r ' << En+aEn_l.
(2.6c)
As in ref. 5) it can be shown t h a t for such a n a r r o w level one has a p p r o x i m a t e l y
tk .12,
(2.v)
so t h a t in this a p p r o x i m a t i o n the c o n t r i b u t i o n of this level to the resonance p a r t of the cross section (1.63) is 4 ~ (2l + 1)
¼ Fz. e
]k,,,J2
(E_E.~)2 +~Fzn2.
(2.8)
F o r E ~ E ~ , this t e r m will be the principal one in the cross section; it is t h e n a p p r o x i m a t e l y equal to the BreitWigner cross section. The same conclusion m a y be r e a c h e d in quite a different way, which has the a d v a n t a g e of showing explicitly t h a t the a s y m m e t r i c t e r m in ./I~ ( E   Ez~ ) in the cross section (1.63) m a i n l y corresponds to the interference between the p o t e n t i a l scattering and the resonance scattering t e r m s appearing in the cross section (1.64). This we shall now see b y proving t h a t ffz. vanishes when the p o t e n t i a l scattering is neglected.
THEORY
OF
NUCLEAR
R E A C T I O N S {I)
545
Let us assume that in a finite range of energy the scattering matrix (1.54), for M = 1, m a y be written in the form
U, = l + k O , ( k ) + k (\kk,,, + k +R,*o k*J'
(2.9)
where the contributions of the resonant states n' :# n have been neglected or incorporated in the potential scattering. From the property (1.28) it is evident that the scattering matrix Uz is zero for k = kn:
k~R*n
1knO'(kn) + ½ R ' n + k,,k*  O.
(2.10)
One has also U,  0 for k k**, but because of eq. (1.29) this leads to the same relation. Eq. (2.10) m a y be used to obtain A~n and B,~, defined b y eq. (1.62), in the form A~ 

2F~~ Im [1ka~)~(k~)],
(2.11a)
Kn
2M ( F ~ Z i m
[1k~)~(k~)]. (2.11b)
When the contribution kaQz(k,,)from the potential scattering is neglected, the corresponding values of A ~ and B ~ simply become those given b y eqs. (2.7). The preceding results also show that if the potential scattering ~)z(k) is known (or neglected), the behaviour of the cross section near a resonance is entirely determined b y two parameters only, e.g. the position E ~ of the maximum and the width F~. Similar approximations could be investigated for a multilevel form of Uz, but the computation of the coefficients Az~,/~z~in terms of the Ezn, Fz~ becomes very tedious. 2.3. ANOTHER EXPANSION OF THE SCATTERING MATRIX Let us return to the expansion (1.54) of U~. One m a y regret, firstly that it is expressed in terms of k rather than in terms of the energy E, and secondly that in each term of the series only the first part Rz~/(kk~ ) is a true term of resonance, while the second part has a modulus which is a monotonic function of k. It is indeed possible to expand U z in terms of E and at the same time to transfer the contribution of these second parts to the potential scattering. For this purpose, let us regard Uz as a function of the complex variable E = ~2k~/2M and let us name the corresponding complex plane the Eplane. The scattering matrix has a branch point at E = 0; to remove this ambiguity we introduce a cut along the negative real axis of the Eplane. Let us make the
546
j.
HUMBLST AND r. ROS~NF~LD
,domain defined b y n ~ arg E ~ +•
(2.12a)
correspond to the half kplane in which   ~1
_< __
arg k =< + fin, 1
(2.12b)
a region which contains the poles k., but not the poles k.*. When there are poles on the imaginary kaxis, these, as well as the corresponding ones on the negative axis of the Eplane, m a y be included in the half kplane considered. ]mk
lm£ B Re~
ReE
8
It,
Fig. 1. The half kplane and t h e corresponding E  p l a n e w h e n there are poles on the imaginary axis. 13 = b o u n d state, V = v i r t u a l s t a t e of d e u t e r o n type.
Fig. 1 shows how this can be done. Two poles on the imaginary kaxis will never be symmetrical about the real kaxis; for if k~  i7~ is a pole of U t , then i7~ will be a zero of Uz. We m a y now use in the cut Eplane expansions of exactly the same form as eq. (1.50); the function e(E) is now only defined as a regular function of E in the cut Eplane. The residue of Ut at the pole St, is simply related to its residue Pz~ at the pole kt= by the factor (?~2/2M)2kt~. Rather than to consider directly an expansion of U z  1, however, it is convenient to start from expansions of the quantities St or St defined by eq. (1.68) : in this way we include in the resulting expansions for the cross sections the factor k 2 occurring in their definitions (1.22), (1.23); the expansions of S t will moreover yield the explicit smallenergy behaviour of the cross sections. For the q u a n t i t y St, the theorem (1.50) gives for M = 0 an expansion of the form r~, S t = g z ( E ) + ~ E_Et,_+_I.
,
(2.13)
where ~2
~t. = ~ k~. R~.;
(2.14a)
THEORY
OF NUCLEAR
REACTIONS
(I)
547
hence
[2ME~½F
]~
7
(2.15a)
E Ez. +~ zFz,~J a t=~r(2lF1) 6,(E)+~
EE,fr'"+ ½ir,.
2.
(2.16a)
Applying the theorem (1.50), for M = 0, to the expansion of St, one gets similarly
U, = 1] (2MEIZ+~"(Q(E)F ~ \ h2 /
r,~
)
(2.15b)
EEzn+½iFz, , '
2,
i
at = n(2lF 1)(2MEIer[
\ h2 ] IQ(E)F~
rzn 1.
EE~ + ~ F ~
(2.16b)
where ~¢~n =
h2
m
rz.
kzn Rzn  k~ln"
(2.14b)
Now, the inequality (2.12a)also implies   a =< arg E* _< n,
(2.12c)
and we therefore still have, in the cut Eplane, a unitarity relation for U~ in the form U~(E*)*Uz(E) = 1. (2.17) We could therefore introduce in the cross section (1.32) the forms (2.15) of Uz which would give for at expansions similar to that given by eq. (1.63). From eq. (2.17), we also derive the property
U~(E,n+½iFzn ) = 0,
(2.18)
which easily leads to the BreitWigner approximation discussed in subsection 2.2. In the cut Eplane there are no relations corresponding to those expressed by eqs. (1.29), (1.28) and (1.56)" such relations would introduce arguments of E which do not satisfy the condition (2.12a). However, one m a y write such relations in an Eplane composed of two Riemann sheets and corresponding to the whole complex kplane. Then one obtains expansions of the following type *:
= O,(k) + z Sln=
~lnPln
F (k+k,.)s,.
~2/2M k~n ,
¢ See Peierls' p a p e r e), p a r t i c u l a r l y p. 21.
(kk*.)s*. 7 + E E l .   l iFz.J '
(2.19)
J . H U M B L E T AND L. R O S E N F E L D
548
where Sz and k are doublevalued quantities. Clearly, this leads for U t  1 to an expansion identical with our expansion (1.54) with M  1. From the foregoing discussion we conclude that in the present case 1) the treatments in a kplane and in a single Eplane with a cut provide different expansions of the scattering matrix, which are equally well adapted to a description of resonance phenomena; 2) the treatments in a kplane and in a double Eplane without a cut provide exactly the same expansions.
3. Elastic Scattering of Charged Particles We now assume that the scattered particles and the target nuclei have the electric charges Zle and Z,e respectively. It is well known that ill order to be able to define outgoing and incoming waves that are singlevalued functions of the complex wave number k, one has to cut off the Coulomb potential at some finite distance r = b. We m a y take this distance b of the order of magnitude of the atomic radius, corresponding to the cut off of the Coulomb field due to the screening effect of the atomic electrons. The radial factor of the partial wave function will then satisfy the equations d* dr2 + k*
[6 2 + k2 Id*
~r#+k 2
/(/+1) r*
v(r)
]
/(/+1)
~]
r2 /(/+1).] uz
u,=O
for
0 _< r < a,
(3.1a)
u~ = 0
for
a <_ r <_ b,
(3.1b)
0
for
r>
(3.1c)
b,
r 2
~,(0, k) = o,
(3.2)
where M = ZIZ2e 2 ~.
(3.3)
The outgoing and incoming radial waves 0~, I~ are solutions of eqs. (3.1b) and (3.1c). For r => b we m a y still define them b y eqs. (1.12) as in the preceding sections. Consequently, all the considerations and results of sections 1 and 2 are still valid in the present case provided that b is everywhere substituted ]or a. This is true in particular of the expressions (1.19) for the amplitudes, the expression (1.20) for the scattering matrix, the definition (1.33) of the resonant states, and their location in the complex kplane. The expansion of the scattering matrix in the form (1.54)is also justified without further investigation. It is not very satisfactory, however, to define the resonant states of a nuclear
THEORY OF NUCLEAR REACTIONS (I)
549
system b y a b o u n d a r y condition at the surface of the atom. We shall now see t h a t the appropriate definition of the outgoing and incoming waves in the interval (a, b) makes possible a description of nuclear scattering in terms of physical quantities taken on the "nuclear surface" r ~ a. Let us introduce in the interval (a, b) two auxiliary Coulomb wave functions /(1), /1(2), defined as solutions of the equation
I~drs2 +
k2
l(/+1) r~"
2;] It u'
= 0
(j = 1, 2)
(3.4)
and satisfying the b o u n d a r y conditions ]l(1)(¢g, k )   I
= /Z(1)'(a, k) = 0,
(3.5a)
/,(2~(a, k)
= f,~s~'(a, k)  1 = 0.
(3.5b)
T h e y are real a n d integral functions of k s for a< r< b a n d consequently have the properties
ly' (r, k) =/y)(r, k),
(3.6)
/,U)(r, k)* =/s(J'(r, k*) /s(J)(r, k*).
(3.7)
T h e y are linearly independent because W ( / (1), /s(2); ~v) = W(/l(1) ' /sis); a) = 1;
(3.8)
therefore, in the interval (a, b) the outgoing and incoming waves 0 z, I s are necessarily linear combinations of such functions. Matching these linear combinations with the radial waves (1.12) at r = b, we i m m e d i a t e l y see that, for a < _ r g b ,
Oz(r, k)  iW(/s (2), ~ 1 ) ; b)/l (1) (r, k ) + iW(]s (1), ~ S ( I ) ; b)/l (s) (r, k), (3.9a)
Is(r, k) : iW(/s~2~,d,~s(2); b)/s(l~(r, k)iW(/~(x),~'C~°s~2); b)/s(S~(r, k).
(3.9b)
But, from the evident relation /y) X~s(~),,_/y),,~fs(k~ 
/z °') ~Vt°t(k~
(3.10)
we deduce
W(/z(J~,~z(~); b) = W(/zu),~,~s(k~; a)2~f:r1/s")~zC~)dr,
(3.11)
and hence Os(a, k) =   i W ( / t ( z ' , w s ( l ' ; b) = i
o's(a, k) =
b) =
[~z(l'(ka)+2r] f2 r1/z(*)3¢~s(l'dr], (3.12a) (k,,)
'1' d,J. ¢3.1 b)
550
Is(a,
J.
H U M B L E T AND L~ R O S E N F E L D
k) = iW(/s{2',..~,{9) ; b) = i EJ~l~tlg~'(kl~)Jf 9.~ f: ,lls{2,...~s(2) d,],
I"(a'k)=iW(/s(l"J/t°s(2';b)=iE
(3.13a)
2/~s(2)'(ka) 2~fabr_ 1Is(l~#f(2~dr , ]. (3.13b)
Defining the function G~(r,r') by
G,(r, r') =/,cl)(r, k) lz(~'(r ', k) /s(1)(r ', k) /,(~'(r, k),
(3.14)
we see t h a t the function (3.9a) m a y also be written as
O,(r,k)  i[~s'l'(kr)+2~T f:r'lGs(r,r')~'l'(kr')dr'~,
(3.15a)
because the function on the righthand side of eq. (3.15a) satisfies the same differential equation as the function (3.9a) and its value and derivative at r = a are also given b y the righthand sides of eqs. (3.12). Similarly one has I s(r) :
  i [ d ~ z (~) (kr) + 2~ j~b r ,_~ (;~(r,r')gff _ .
,(~(kr')dr'].
(3.15b)
F r o m the properties (3.6) a n d (3.7) of the functions ]s (j) it is evident t h a t the functions 0s, Is have the following properties for a <_ r ~< b as well as for r > b: = L(k),
Os(k ) = (1)*Is(k),
oak),
L(k*)* = z,(k)
= (1)'o,(k),
(3.16a) (3.16b)
whence also
Oz(k* ) = (1)sO~(k) *,
5(k*)
= (  1 ) s z , ( k ) *.
(3.17)
In practical calculations, the functions ]s (1), [s (~ introduced here will be expressed as linear combinations of the usual Coulomb wave functions, i.e., the W h i t t a k e r functions corresponding to outgoing and incoming waves, or the regular and irregular functions usually denoted Fs, Gz. The matching of such linear combinations with ]s (1) and ](2) is easily m a d e at r = a, as a consec uence of the conditions (3.5). Further, we still have
W(O s, Is; r, k)  2ik.
(3.18)
The functions Os, Is, u, being solutions of the same equation in the interval a <_ r _< b, the Wronskians W(us, 0~; r) and W(u,, Is; r) are also independent of r, a n d we have in particular
W(u t, Os; a) = W(us, 0,; b),
W(u s, Is; a)  W(u s, Is; b).
(3.19)
These relations have the following consequences: all the results of the preceding sections, such as the expression for the collision m a t r i x
W(uz, I d a, k) Us = W(us ' Oz; a, k)
(3.20)
Tn=o~,
oF
N~CLSAR
R=ACT~O~S
(~)
551
and the definition of the resonant states
W ( u t, 0~; a, k) = 0
(3.21)
m a y be applied to the present case without substitution o / b / o r a provided that the expressions (3.15) for the outgoing and incoming radial waves are used instead of those given b y eqs. (1.12). For the discussion of the properties of the resonant states it is important to observe that the property expressed b y eq. (1.42) holds in the present case, whether or not a is replaced b y b. In fact, if we define
O't(a, k) L,(k)  O,(a, k~'
(3.22)
we have 1
Lt(k)L,(k)*  10,1~ W(O,*, 0~, a, k), or, in virtue of the identity
w (o,* , O,; b, k )  w (o,* , o,;
k) =
(07 O,*O,*" O,) d,, :
and the fact that
O,(b) ~ i~°z(l'(b),
O't(b) ~ iW,~"(b),
to,l' dr
(3.23)
the equivalent expression
IO,(a)i2[L,(k)Lt(k) *] = W ( ~ (1,*, ~taz(l'; b, k)~(k2k *2) fba [O,(r)i2dr. With the help of eq. (1.42), in which a is replaced b y b, this indeed yields a relation of the same form,
lOt(a, kn)l~[L,(kn)L,(k,)*] = i(kn÷kn*)lWoCl'(knb)l ' +(k~2kn *~) [ f b a IOt(r,k~)12dr+.A/'t.(b)~.
(3.24)
The same argument as in subsection 1.3 leads for ~ , when K~ =# 0, to the essentially positive expression 2y.
:o( do ]utn(r)l'dr÷
[u,~(a)/O,(a,
kn)121ei~b[ ~
(3.25)
O,(a, kn)
In the expression (3.25), however, the exponential factor exp (2~,b) in the numerator seems to give the physically irrelevant parameter b undue influence on the value of ~,: actually, of course, this factor is compensated b y the similar increase with b of the denominator. It is therefore advisable to transform the formula (3.25) so as to remove the explicit occurrence of the factor in question.
552
J . H U M B L E T AND L. R O S E N F E L D
This is readily achieved b y observing t h a t the decomposition of the righthand side of eq. (3.24) into two terms respectively proportional to
i(kn+kn*) = 2iKn
k~2k, .2 = 2iK~ • 2y~
and
is obviously not unique; thus, b y writing ]j/,~o~l~(k.b)]2
=
1+27 n
leik~b]2 =
fb [e~k.r]2dr' 0
(3.26)
we get a new decomposition of the same form, from which the terms exhibiting the u n w a n t e d increase with b have disappeared:
IOn(a, k~)12[Lz(k~)Lz(kn) *] = 2iJ%E12~nN~n(a)] ,
(3.27)
with
This gives for ~n the new formula, of the same t y p e as (3.25), 2yn =
lu,.(a)/O,(a, k.)l 2
a
f
0 lu'"(r)12dr+
0,(a, u,.(a)
;
(3.29)
2 N,.(a)
it is no longer assured t h a t the coefficient N~, (a) is nonnegative, b u t it is clear t h a t the d e n o m i n a t o r of the expression (3.29) is still positive. In other words, we use the decomposition (3.24) to prove t h a t y, is positive, and we conclude from this result t h a t the equivalent expression (3.29) with the value (3.28) of t h e coefficient Nz, (a) also has the property of being the ratio of two essentially positive quantities. In conclusion, we m a y mention explicitly t h a t the "aindependence p r o p e r t y " of the resonant states kn still holds. In fact, for a n y x :> a, the solutions of the equation
W ( u , , 0,; x, k) = O
(3.30)
are independent of x because u s and O~ are two solutions of the same equation for such values of x. Of course this presupposes t h a t the potential remains the same for different choices of x: it m u s t be v(r) for r < a, 2~/r for a _< r < b and zero for r > b.
Part II: THE GENERAL CASE OF NUCLEAR R E A C T I O N S 4. The Configuration Space 4.1. GENERAL
ASSUMPTIONS
AND
NOTATIONS
Our general assumptions will be the same as those made by Lane and Thomas 2) in their review article on the Rmatrix formalism: conservation of
THEORY
OF NUCLEAR
REACTIONS
(I)
553
probability, time reversibility and causality in the framework of nonrelativistic mechanics. Whenever possible, we shall adopt the same notations as Lane and Thomas 2) in order to facilitate reference to their paper if more details are wanted about the configuration space, the elimination of the barycentre motion, the channelspin wave functions, the wave functions of the relative motion, and the definition and properties of the collision matrix. These topics, which are of course independent of the Rmatrix formalism, are discussed only briefly in the following. The discussion of channel wavefunctions and boundary conditions in this section is in all essentials also the same as that adopted b y C. Bloch 16). 4.2. REACTION CHANNELS We shall call channel a situation in which the total system of A nucleons we are considering is separated into two fragments, containing A 1 and A s nucleons respectively, each of which is in a definite stationary state, while their state of relative motion is also specified. For simplicity, we shall confine the discussion to reactions involving only channels of this restricted type, i.e., reactions in which only two nuclear systems interact to form two other such systems, without a n y creation or annihilation of particles; photons are also excluded from the discussion. The state of each fragment A t is characterized b y a total spin quantum number I~, its projection ij on some fixed direction in space, and a set of other q u a n t u m numbers specifying the energy of the state, which we denote by ~ . Instead of il, is, we m a y also use the quantum numbers s, v specifying the channel spin s  I i + I s ,
]IiIsl
< s < 1 1 + 1 3
(4.1)
and its projection v on the fixed axis. We shall consider states of relative motion of definite orbital m o m e n t u m 1, with quantum numbers l, m. Because of the conservation of the total angular momentum J = l + s , of quantum numbers J , M, we m a y replace the q u a n t u m numbers v, m b y J , M in the definition of the channel. We shall thus understand the channel symbol c to represent a n y one of the sets of quantum numbers
{o~1o~2(IlI2)ili2, lm},
{o~1o~2(IlI2)sv,lm},
{~l~S(IlI,)sl, JM}.
For each mode of subdivision into two fragments A1, A2, we use an appropriate set of coordinates, to which we attach an index chosen as the same letter :~ by which we denote the states of excitation ~1, :c~ of the fragments. Thus, we shall take for the two fragments sets of internal coordinates q~l' q~, (including the spin coordinates of the individual nucleons), and relative position coordi
554
I.
H U M B L E T A N D L. R O S E N F E L D
hates r~ of their barycentres. The motion of the barycentre of the total system is, as usual, eliminated. The possibility of treating the two fragments as separated results from the limited range of the nuclear forces; in other words, if r~ denotes the distance between the barycentres of the two fragments, there exists a distance a~, the "channel radius", such that for r~ > a~ only nonpolarizing forces (in fact, only Coulomb forces) act between the fragments. As in the preceding section, we shall cut off the Coulomb potential at some distance b~ of the order of atomic dimensions. The radii a~ and b~ are, of course, introduced only for mathematical convenience, and will not be allowed to play any essential physical part in our theory. 4.3. I N T E R I O R
AND EXTERIOR
REGIONS
The positions of the A nucleons can be represented b y a single point in a configuration space of 3A dimensions. The region of this space corresponding to the A nucleons being close together in physical space is called the interior region; more precisely, under the simplified conditions adopted here (ignoring the possible splitting into more than two fragments), the interior region is defined as the part of the configuration space in which r~ < a~
for all c¢.
(4.2)
The boundary sur/ace of the interior region is then composed of an assembly of parts of the surfaces r~ = a~. Any point which is not in the interior region or on its boundary surface ~ce is said to be in the exterior region. To be really useful, this distinction of interior and exterior regions must be applied to a configuration space in which one more coordinate, the spin coordinate, is added to the space coordinates of each nucleon. The interior region is then still defined b y the conditions (4.2). Let ~ denote the portion of the surface r~ = a~ which is part of the boundary surface ~9°. In this sense we think of the surface 5~' as composed of partial surfaces ~9~ such that a point of the configuration space lying in the exterior region just outside ~ga~ corresponds to the subdivision e of the A nucleons. We shall then have c~ = ~ , and an element of the surface 5z~ will be defined b y
dcga~ = a~2 d~Q~dq~ldq~z ,
(4.3)
where dD~ is the element of solid angle corresponding to the relative angular position of the pair a, and where dq~ corresponds to the internal coordinates of the fragment Aj. A symbolic integration over d~9°~ implies not only an integration over the internal space coordinates, but also a summation over the spin coordinates of the individual nucleons.
THEORY OF NUCLEAR REACTIONS (I)
555
5. T h e W a v e  f u n c t i o n in the E x t e r i o r R e g i o n 5.1. DECOMPOSITION
OF THE WAVEFUNCTION
IN RADIAL
AND SURFACE
FACTORS
In a channel c, the total wavefunction kve m a y in the first instance be analysed into two factors, pertaining respectively to the internal state of the fragments and to their relative motion. Owing to the spin degeneracy in the external region, where the fragments only interact through the Coulomb potential, the internal wavefunctions W~1I1~, ~ O a , i , i , of the fragments m a y be grouped into a channelspin wavefunction
~oss~ =
~.
(Ilil,
I2i2[s~,)v,~,j~q v'~,z,,,
(5.1)
ix+i 2 = v
completely characterizing the internal state of the system; in this and later formulae, symbols like (Ilil, I2i2]sv ) denote ClebschGordan coefficients and & is written for ~la,, while indices 1112 are generally omitted. The states of relative motion of the two fragments m a y be represented b y products of spherical harmonics ilY,~t(~2~) and radial components uc(r~)/r~. It is convenient to single out the radial factor and to group together the other two into a "surface factor" embodying the dependence of the wavefunction on all coordinates except the relative distance r~. For this surface factor we m a y write, in the representation {asv, lm},
Ym t ~°~sv,
(5.2a)
(SV,lm[JM)qDa,,,,m.
(5.2b)
q)ssv, ~m =
i z
or in the representation {asl, JM}, ~P~*Z,IM =
]~ v.kin= M
For all channels corresponding to the mode of fragmentation a, the surface functions 9e m a y be chosen to form an orthonormal set of the usual kind, which m a y be assumed to be complete in the subspace of all coordinates except r~. In this subspace, we write the orthonormality relation in the form f ~e*~e' dS~ = 8ce',
(5.3)
with dS~
d5¢~ ~a 2  d~9~dq~ldq~,.
(5.4)
The completeness of the set 9e allows us to express the exterior wavefunction !gext in its most general form, at any point of the exterior region corresponding to the fragmentation x, as a superposition of all the channel
556
J.
HUMBLET
AND L. ROSENFELD
wavefunctions ~e belonging to this mode of fragmentation :~: ~ext,a, = X ~T/c= X (~c Ue(~'a), c(a) c(a) ra
(5.5)
where the symbol ~c(=) indicates a summation over all channels for which the mode of fragmentation of the compound system is a. Moreover, since the very existence of the channels implies that the overlapping of two functions Text(,~ , ~JexU=') is completely negligible, the general form of the total wave function in the exterior region is = X
(5.6)
or
= X c
= Z vo c
/'a
,"
(5.7)
in eq. (5.6), only one term can be different from zero at some specified point of the exterior region, while the sums in eqs. (5.7) are extended to all possible channels, irrespectively of the fragmentation. 5.2. WAVE NUMBERS AND RADIAL FACTORS FOR OPEN AND CLOSED CHANNELS The total energy o~ of the compound system is expressed in a given channel c as a sum of the internal energy E~E~I+E~,, of the fragments and the energy Ee of their relative motion:
@ = Ee+Ez.
(5.8)
If Ee > 0, the channel c is said to be open; if Ee < 0, it is said to be closed. The relative motion in an open channel is characterized by a real positive wave number defined by ,
(5.9)
where M~ = M~IM~,I(M~I + M~,) is the reduced mass of the fragments. The corresponding relative velocity is ~kc vc  M~"
(5.10)
In a closed channel, the wave number defined by eq. (5.9) is imaginary; it will be taken with Im k e > 0. (5.11) For a state of the system of given energy o~, the various channel wavenumbers are not independent, but axe connected b y the relations of the form (5.8)
THEORY OF NUCLEAR REACTIONS (I)
557
expressing the conservation of energy in the processes concerned:
 2M~ ke*+ E~  2M~ k~c'+ E i . . . . . .
(5.12)
The radial factor uc(r~) of the channel wave function ~Vc appearing in eq. (5.7) m a y be written, as in the singlechannel case treated in the first part,
ue(r~) = xcO,(r~, ko) + yelz(r~, kc),
(5.13)
where the functions Oz, I~ are again those defined b y eqs. (3.15), with the value of the orbital momentum l belonging to the channel c. In this and the following two sections, we consider only the positive and the purely imaginary values of the wavenumbers ke, which have the direct physical meaning just explained. In sections 8 to 10, when we come to discuss the expansion of the collision matrix, we shall again introduce complex parameters kc, as we did in part I. 5.3. T I M E
REVERSAL
PROPERTIES
If the usual phase convention for the spherical harmonics is adopted and the phases of the internal wave functions Y~,v are appropriately fixed z), it is easy to see that we have
Kg~a, JM = (1) JM 9~a,JM.
(5.14)
Since the operator K acts on the radial factors 0z, I z as the complexconjugation operator, we simply have, using the relations (3.16),
KOz(kc) =I~(kc*),
KIz(ke ) = Oz(kc*).
(5.15)
Hence, for an open channel, we have KO~(kc) = I~(kc),
KI,(kc)  O~(k¢),
(5.16a)
while, for a closed channel,
KO,(ke) = Iz(ko) = (1)'0~(kc), KIz(ke) = O,(ke) = (1)'Iz(ke).
(5.16b)
6. Boundary Behaviour of the Interior Wavefunction 6.1. THE INTERIOR RADIAL FACTORS In the case of elastic scattering by a fixed potential, the interior channel wavefunctions can be factorized in the same way as the exterior ones, and the boundary conditions can accordingly be expressed in terms of the radial factors only. In the general case, no such factorization of the interior wavefunction exists, because of the couplings between the channels, due to the interaction of
558
J.
HUMBLET
A N D L. R O S E N F E L D
all the nucleons in the interior region. Nevertheless, on the boundary surface itself, quantities playing the same part as the interior radial functions u,(a) and their derivatives u',(a) in the simple case can be defined b y taking advantage of the following facts: 1) although the "radial" channel coordinates r~ lose their physical meaning in the interior region, they are still mathematically defined in this region; 2) the boundary surface can be decomposed into adjacent parts ~ corresponding to the various modes of subdivision of the compound system, and 3) on each such part ~ , we m a y expand any function with the help of the associated orthonormal set of surface functions ~c. In particular, at any point of 5a~ we can give a definite meaning to the value of the interior wave function Tin t, multiplied for convenience b y r~, and to the derivative with respect to r~ of this product r~Tlnt: we simply have to give the coordinates in which r~kUint and (~/~r~)(r~Tint) are expressed values satisfying the condition r~ an. We m a y then expand the surface functions so defined in the form
(r~Wtnt),~=%  Z ¢cq~c,
(6.1a)
c(a)
(6.1b) rot = a a
c (a)
where the expansion coefficients are given b y
Ce = f ~o* (r~ T~t)~,= ~ dS~, ~b'c ~
¢pc* ~r~ (r~Tint)
(6.2a)
dS~.
(6.2b)
r ~ ~ aGt
It is clear that the quantities Pc, O'e reduce to the radial factors u,(a), u',(a) in the singlechannel case, and we shall see that they form the natural extension of these factors to the general case. In fact, we shall call qs' e the "derivative" of qSc and speak of an expression like qgcv ~' e q~'O c e as the "Wronskian" of q~c and 0e at r~  a~, which we shall denote b y W(¢e, 0e). Because we neglect the overlapping o f two functions q~e corresponding to different fragmentations, we m a y also give to the expansions (6.1) the following form: (rTlnt)~ = ~ q~eq~e,
(6.3a)
c
X e~Ve, I~r (rV',nt) 1 :  c qY
(6.3b)
where the lefthand side is taken at any point of ~a, while in the righthand side
THEORY OF N U C L E A R REACTIONS (I/
559
the summation is now extended to all the channels, independently of the fragmentation to which they belong. 6.2. G R E E N ' S
THEOREM
Before discussing the boundary conditions, we shall point out, for later use, a general relation between the product of any two eigenfunctions of different (real) energies 6°, integrated over the interior region, and a channel sum of Wronskians of the corresponding radial factors defined b y eqs. (6.2.). This relation follows from a straightforward application of Green's theorem to the integral of a suitably chosen expression over the interior region. Let us consider two solutions T 1, T2 of the wave equation H~Poz~P, corresponding to two different real energies 6°1, oz2. Then, we m a y write (~2°~1) f~ T2* Tide) fo) [(H}tr/2)*~rYl}[Y2*(H}//1)] d°),
(6.4)
where the integration in configuration space is extended to the interior region co bounded b y the surface 50. If we assume the potential V to be selfadjoint, i.e.,
f~, [(VT2)*T1T~*(VT1)Jdco=
0,
(6.5)
only the terms of H corresponding to the kinetic energies contribute to the righthand side of (6.4), and these are transformed b y Green's theorem into an integral over the surface 50. For the part 50~ of the surface, this integral can be written as
fr 2M~ J L
. ar~
Tlr
*
a
T1)l
, dS~, ~r~ _j r~= %
or, on account of the expansions (6.1) and the orthonormality relations between the ~e,
1i3
t,
,
r
2M~ c(~) The integral over the whole surface 5 ° is the sum of similar expressions for the different modes of fragmentation :¢, and eq. (6,4) thus becomes
(~2d~1)f
T2* T l d c ° =
~c 2Me W@*o, ¢1o),
(6.6)
where the summation now extends to all channels, and where we have written Me instead of M~ for the reduced mass in channel c. Eq. (6.6) is the announced general relation, from which we shall derive those we need b y an appropriate choice of ~1 and ~ .
560
j.
H ~ M B L ~ T A N D L. R O S ~ N F E r O
6.3. T H E B O U N D A R Y CONDITIONS
With the help of the representation (6.1) of the interior wavefunction and its normal derivative on the partial boundary surface 5 P , we immediately obtain, b y comparison with the expansion (5.5) of the exterior wavefunction, the continuity conditions on this surface. Because of the linear independence of the 9e, these conditions become, with the expression (5.13) for uo, ~be = xe 0 t (a~, kc) +yelt (a~, ke),
(6.7a)
#'e = xe0't(a,, ke)+yeI'z(a,, ke),
(6.7b)
for all channels c belonging to the mode of subdivision x. Repeating the argument for all the subdivisions, we thus arrive at a complete formulation of the boundary conditions for all channels, which has the advantage of yielding the amplitudes Ze, Ye in explicit form for each channel separately: 1
xe yo = +
2ike W(~e, It; a~). 1
w @ c , or; a,).
(6.8a)
(6.8b)
It is through the factors #e, #'o that a coupling is implicitly established between the amplitudes in the various channels, inasmuch as specification of the incoming amplitudes fixes the interior wave function and all the radial factors ~o, #'o which then, b y eqs. (6.8a), determine the outgoing amplitudes for all channels. From now on, when no confusion is to be feared, we shall often find it convenient to use the channel index c instead of more specific indices such as l or ,¢ for quantities like 0~, It, a~. 6.4. W A V E  F U N C T I O N S W I T H A S I N G L E E N T R A N C E CHANNEL
It is easy to reduce the problem still further b y noticing that any interior wavefunction specified b y any number of incoming amplitudes ye (which in practice will all belong, of course, to the same mode of fragmentation x) can be expressed as a linear combination ~rJ ~__ X ~r/(e)
(6.9)
c
of wavefunctions ~ c ) corresponding to a single entrance channel c, i.e. for which Yc' = 0 for c' ¢ c, (6.10) or, more explicitly, according to eqs. (6.8b),
W(q~Cccl, Oc,) = 0
for c' ~ c.
(6.11)
THEORY"
OF
NUCLEAR
REACTIONS
(i)
561
The upper index (c~ affecting the various quantities serves to remind us t h a t t h e y refer to the case in which there is no other entrance channel t h a n c. According to eq. (6.9) one also has ¢c' :
t
~ ~b~cC),
¢c' : Z ~c'(n(c)',
c
(6.12)
C
and therefore, on account of eqs. (6.11), oc,) = X
=
(6.1s)
C
which means, according to (6.8b), vc, =
,
(6.14a)
while on the other hand, b y (6.8a), x c, = • xc
(6.14b)
c
We therefore see t h a t it is sufficient to treat the case of a single entrance channel. We m a y assume that, except for a normalization factor, the solution ~(c) of the wave equation is completely defined b y the eqs. (6.11)   which m a y be regarded as b o u n d a r y conditions on ~¢   and the condition of finite probability: [T(C~l2 finite in the interior region. (6.15) The latter condition is a generalization of eq. (1.2b) for the elastic scattering; it excludes in particular a n y wave function whose a s y m p t o t i c behaviour in channel c would be t h a t of a purely outgoing wave, with no incoming wave in this channel. In the representation (~sl, ]M}, it is clear that, because of the conservation properties of J and M, we have ¢c(e,) = q~
when
J , M 5~ J ' , M ' .
(6.16)
Moreover, since ~(aa, JM)' ~'8' v,JM) IM' ~(~81, ~',' V YU are scalar products, t h e y are independent of M. In particular, if the n o t a t i o n c is used to designate the channel {Ksl, JM}, we have (c)
at
=
~ (_~) .
(6.17)
7. T h e C o l l i s i o n M a t r i x 7.1. D E F I N I T I O N
Let us return to the general state described b y the eigenfunction (6.9). For the component ~(e~ corresponding to the single entrance channel c we
562
j. HU~BLST ^ND L. ROSSSrELD
m a y write, according to eqs. (6.8), X(c¢,) =  Uc, e Y(cc)
(7.1)
with
ue,¢ =
W (kq ~ccl C t ' Ic,)/ke, oo)/kc
(7.2)
From eqs. (6.14) we then derive the relation
xo, =  X
uo, o vo,
(7.3)
c
showing how in the general case when there are several entrance channels, with given amplitudes ye, the amplitudes of all exit channels are determined as linear combinations of these incoming amplitudes. The coefficients Ue'e of the linear relations (7.3) form a matrix !J which is fundamental for the computation of the crosssections for the reactions leading from any one of the entrance channels c to any exit channel c'. Hitherto no distinction has been made between open and closed channels, and the definition of the matrix U is perfectly general in this respect; but of course, only those matrix elements Ue,¢ which refer to open channels c', c, i.e. to real and positive values of the channel wave numbers ke, k¢,, are susceptible to a physical interpretation in terms of a reaction leading from channel c to channel c'. If the wave numbers ke, ke, are real and positive, and ve, re, represent the channel velocities according to the definition (5.10), the incoming and outgoing fluxes in the reaction c  + c' are given b y vc{ycL 2 and Vc,lXe,]~; it is therefore natural to define the element a//e,c of the collision matrix ~ b y
=
Uo, e;
(7.4)
the cross section for the reaction c  + c' will then be directly expressed in terms of the amplitude q/c'e Yc(un), where yc (un) is the amplitude corresponding to unit incoming flux. In the following, however, we shall also call U "collision matrix" when no confusion is to be feared. The definition (7.2) of the matrix element Ue, e shows that it is entirely determined b y the consideration of the wavefunction with the single entrance channel c, which was introduced in subsection 6.4. The most convenient channel designation scheme for representing the matrix U is obviously c ~ {asl, J M } : in this scheme, according to eqs. (6.16), the matrix is diagonal with respect to the angular momentum quantum numbers J, M. Moreover, its elements (like the radial factors q~,) on which they depend) are independent of M and m a y therefore be written c,o =
s
(7.5)
THEORY OF NUCLEAR REACTIONS (I)
to the scheme
{~sv, lm},
(s'v', l'm']JM)(sv, lm]JM)U ]~,.,,z. ~l •
(7.6)
One easily passes from the r e p r e s e n t a t i o n for instance, b y relations of the form 2) Ui,., v,. ~,m,;~,.z~  ~ 7,2. G E N E R A L
{asl, JM}
563
PROPERTIES
F r o m their v e r y definition, it is clear t h a t the collision m a t r i x elements
Uqp are i n d e p e n d e n t of a n y p a r t i c u l a r n o r m a l i z a t i o n a d o p t e d for the solution y]~m of the wave equation. T h e general properties of the collision m a t r i x are most easily established b y considering a wave function ~ m for which the n o r m a l i z a t i o n yp = y~ppl= 1 has been adopted. According to eq. (7.1), this yields the following a s y m p t o t i c b e h a v i o u r of ~ m in the exterior region:
~,,~, : ~ Ip  Xc vo~q,o °° = X¢ ~° (~o~Iovo~oo),
(7.7)
whence for its radial factors the expressions
• ~p~ = ~pxe(,o, k ~ )  u ~ o ~ ( a o , ko).
(7.8)
A similar n o r m a l i z a t i o n will be a d o p t e d in this subsection for all the wave functions with a single entrance channel we are going to consider. 7.2.1.
Unitarity
L e t us consider two solutions ~v1 = hv~p~, ~2 = ~cq~ of the wave e q u a t i o n corresponding to the same e n e r g y 0*1 = @2 = @, for which the channels p, q are b o t h open. According to Green's t h e o r e m (6.6), we h a v e
0, where the sum is e x t e n d e d to closed as well as to open channels, which will be respectively d e n o t e d b y c  a n d c +. Eq. (7.8) is valid for b o t h c = c + and c = c, while from eqs. (3.16) we h a v e
qbtcq)* = ¢ScqOc(kc)U*qlc(kc) ~{cql*  (1)z,[Ocqlc(kc)U*qOe(kc)]
for c ~ c +,
(7.10a)
for cc.
(7.105)
I n t r o d u c i n g the expressions (7.10) into eq. (7.9), a n d taking eq. (3.18) into account, one obtains
X
?/2
(2ikce}cpf}cq~2ikcV*cqUep) ~ (2iko~eqU.p+2ik.~.,U*q)(1)'o +Y.~ C
= o.
564
J. HUMBLET AND L. ROSENFELD
All the terms of the latter sum vanish simply because p and q are open channels, and we are left with the identity c~.~kc
u~%uo~= ~?tkp ~..
(711)
Using the definition (5.10) of the channel velocity, this m a y be written (7.12)
vo Uoq U~p = % ~pq, c+
or, in terms of the collision matrix (7.4),
X ~cq°Yep * = 6pq. c+
(7.13)
Eq. (7.13) expresses the unitarity of the collision matrix ~d in the subspace of open channels: q/* ad = 1. Eq. (7.11) yields, for p  q,
(7.14)
F. vclVopl 2 = vp, C+
expressing that the total outgoing flux, for all reactions initiated in channel p, is equal to the incoming flux. 7.2.2. Symmetry Let us now adopt the {~sl, J ~ channel designation and let us again use the notation q to designate a channel identical with q, except for the opposite sign of the angular momentum. If T(q) is an eigenfunction of the Hamiltonian characterized by the single open entrance channel q, two other eigenfunctions are
T(q) = ~~fpq~P(O p_ql pU p_qO_p)
(7.15)
and
( 1)],+M.K~(")
= X ~_2 [6pqOp (kp*)   V~q Ip (kp*)] p
Yp
= X ~" [ 6 ~ , o ~ ( k , )  u ~ J , ( k , ) ]  X p+ r p
p
u~,o~(k,)il)'.; (7.16)
to derive the last formula, use has been made of the properties (5.14) and (5.16), and account has been taken of the fact that the channel q is open. The wavefunction (7.16) is characterized b y amplitudes U~Q in all open entrance channels p (and zero amplitude in the closed entrance channels); accordingly, it m a y be represented as a linear combination   ~ c + U* ~ c q W(c) of single entrance channel functions of the form (7.7), viz. ~ q'P UI,o U~qOp

c + p+ r p
 Upc U~q Op. ~' Upq/p + ~ ~ (PP ¢+ p  ~'p rp
(7.17)
THEORY OF NUCLEAR REACTIONS (I)
565
The comparison of the equivalent expressions (7.16) identification of the coefficients of 0p+, that
and (7.17) shows, b y
Upc Vcq = ~pq,
(7.18)
C+
when p and q are two open channels. In the subspace of open channels, this m a y be written in the form U U * = ~ * = 1. Combining eqs. (7.12) and (7.18) gives for any pair of open chalmels Vp V p q =
Vq V q p ,
(7.19)
a'~qp,
(7.20)
i.e. a'~pq =
a relation showing that the collision matrix is symmetrical.
8. Complex W a v e N u m b e r s 8.1. T I M E
REVERSAL
PROPERTIES
With a view to introducing a definition of the resonant states and an expansion of the collision matrix elements along the same lines as in part I,
we shall from now on consider all the channel wave numbers kc, kc,, . . . as complex quantities. In this subsection we provisionally treat them as independent, ignoring the energy conservation relations. The definition of the functions ~(c) is easily extended to this case: it is a solution of the wave equation
ke2+E~ T (c),
H T (c) =
(8.1)
in which the energy is expressed in the form corresponding to the selected entrance channel, and it is completely characterized b y the conditions
Oo,(ko,))
= 0
for c'
c,
(8.2a)
1~¢)[2 finite in the interior region,
(8.2b)
as a function not only of the space and spin coordinates, but also of all the complex wave numbers. The choice of the normalization factor for the function Tee) (kc, k c , , . . . ) is restricted b y the specification of its timereversal behaviour, which has to be compatible with the boundary conditions (8.2). In the {&sl,JM} channel designation, such a requirement is fulfilled, as we shall see, b y the specification
K~(e)(ko, ke',
• • ") =
~( I ~( ]  M   ~rl(e~ k J ~
o * ,   k * e ' , • • ")"
(8.3)
566
J.
HUMBLET
AND
L.
ROSENFELD
For the timereversal transformation of the radial factor ~5c¢c,) (with J = .]',
M=M'), K~bc¢C)(ke, k e , , . . . ) K t
= f (Kge,)*(r~K~C~t),~=a dS~,
we have, according to eqs. (5.14), (6.17) and (8.3), ~b(cc,)(kc, ke, I
•
•
.)* = (bIel ~ C t (4
* ,ko,,*
• •
.).
(8.4a)
Similarly, we have for the derivatives ¢¢cC)'(kc, ke', • • ")* = ¢¢e)'tc',ke*, kc* , .. .).
(8.4b)
Together with eqs. (3.17), these equations (8.4) show that the b o u n d a r y conditions (8.2a), written for the set of wavenumbers k e* ,   k c*, , . . . are just the complex conjugates of the same conditions written for the set ke, ke,, . . . . This establishes the required compatibility of the time reversal behaviour (8.3) of the function T (e> and the b o u n d a r y conditions defining this function. In the same framework of complex independent wave numbers, if the amplitudes c'~(c>, yCcC)and the collision matrix element Ue, e are regarded as defined b y eqs. (6.8) and (7.2) in terms of the radial factors ~b¢cC,)(kc, ke,, • • ), it is clear from eqs. (3.17) t h a t t h e y satisfy the relations t
Kx~cC)K¢
XCcC)(kcke,,
.),
(8.5a)
KyCcc)K* Ye"(c) (ke, ke,, .. .)*  ( 1)~y¢cC)(ke*, kc*, .. .),
(8.5b)
,
" "
.)*
~
tk   l~'x(C)t k e * ke*, ] c' k,
" "
KUe, cg¢ = Uc, c(kc, k e , , . . . ) *  (   1 ) "  ' U c , c(kc*, kc* . . . . ).
(8.6)
8.2. T H E P H Y S I C A L koPLANE
Let us now impose a restriction upon the range of complex values which we allow for the channel wavenumbers b y extending to the complex domain the validity of the energy conservation relations (5.12); the internal energies E~ . . . . keep their physical real values. We m a y then consider all the channel wavenumbers ke, kc . . . . to be functions of one of them, which we conveniently choose to be the w a v e  n u m b e r k o of the channel of lowest internal energy E~0, i.e. of highest kinetic energy h2ko2/2Mo . E v e r y w a v e  n u m b e r ke is given in terms of k o b y an expression of the form k c ~= k c (k0) = %/k02Kc 2,
(8.7)
where Kc is a real q u a n t i t y defined b y K e 2 = E~E~o ,
(8.8)
and which we shall take to be positive. ? I n the article of Lane a n d T h o m a s 2) sign factors (1) * are missing in t h e i r eqs. (VI.2.12), w i t h the result t h a t the sign factor is also a b s e n t f r o m their relation (VI.2.20), which should be the s a m e as o u r eq. (8.6).
THEORY OF NUCLEAR REACTIONS (I)
567
The values +Kc, Ke of the channel wavenumber k0 are branchpoints of the function ko(ko) defined b y eq. (8.7). To make this function singlevalued we must introduce into the complex k0plane an appropriate cut joining the two branchpoints. We shall draw this cut (fig. 2) below the real axis and symmetrical with respect to the imaginary axis and choose the branch of the function ke(ko) which is positive when k0 is real and positive and larger than Ke.
~, I
K~ i
J
T
 ael~
Fig. 2. T h e p h y s i c a l koplane w i t h t y p i c a l c u t s b e t w e e n c h a n n e l t h r e s h o l d s .
This ensures that we have in general the important property
ko(k0*) = ke(k0)*,
(s.9)
and that the critical values ¢Kc have a simple physical meaning: according as the (real and positive) wavenumber k0 is larger or smaller than Ko, the wavenumber kc is real and positive or purely imaginary with positive imaginary part, which corresponds to the channel c being open or closed; the point Kc refers to the timereversed situation. We m a y therefore call the points ± K e the thresholds of the channel c in the complex k 0 plane. The domain obtained b y excluding from this plane all the cuts joining the thresholds will be called the
t~hysical koplane. It is clear that the formulae of sections 5 and 6 are immediately extended to complex wavenumbers and energies defined in the physical koplane. It must be pointed out, however, that the important formula (6.6) resulting from
568
J.
HUMBLET
A N D L. R O S E N F ] ~ L D
Green's theorem has to be slightly modified: it holds in the form
(G*el) fo 2* ldo
= Zc
(8.10)
always provided that the hermiticity condition (6.5) is fulfilled. Thanks to the relations (8.9), the results of subsection 8.1. also remain valid in the physical koplane. Instead of the relations (8.3), (8.4) and (8.6), we might now simply write g ~ ( c ) (ko) :
( 1)3"M~°'(c) (ko*),
• (:,)(ko)* = c'
~'c' (k0)* = ¢(cc)'(k0*),
uc, o(ko)* =
(8.11) (8.12)
(8.13)
In practice, however, it is also useful to keep the explicit notation (ke, kc,, • • .) of subsection 8.1 in the present case of nonindependent complex channel wavenumbers. In particular, for real values of k o ( ~ qKe, 4Ke,,...), the relations (8.12) and (8.13) are more clearly written in the form ~b(C)c'~/k+, k_)*  c'~(c)(k+, + k _ ) ,
• (d)'(k+, k ) * =
+k_),
G,c(k+, k ) * =
o(k+, +k_),
(8.14) (8.15)
where k+ refers to all the open channels, k_ to all the closed channels. It will be noticed that eqs. (1.25b) and (1.29) are special cases of eqs. (8.12) and (8.13). 9. T h e R e s o n a n t S t a t e s
9.1. DEFINITION In this section, we define the resonant states, in direct extension of the procedure of section 1, b y boundary conditions expressing the absence of any incoming wave. We formulate these conditions in terms of the channel wavenumbers defined in the physical koplane, and adopt the {&sl, J M ) channel designation. Because of the conservation properties of the angular momentum, we m a y discuss separately the states of our nuclear system corresponding to each given set of quantum numbers J , M: in what follows, all channels are understood to correspond to the same set J , M. We now write down simply the conditions expressing the absence of incoming waves ill any channel b y arbitrarily selecting some channel c and imposing upon the radial factor ~b(¢clof the state with a single entrance channel c the condition
W¢~ (c) Oc) = O, k C
(9.1a)
THEORY
OF NUCLEAR
REACTIONS
(I)
569
which, according to (6.8b), indeed means Yc = 0. Besides eq. (9. la), the eigenfunction T(e) under consideration m u s t satisfy the relations W(~(ce), Oe, )  0
for c' ~ c,
(9.1b)
expressing t h a t c is the only entrance channel. The conditions (9. la) a n d (9. lb) together form a set W(~Oe, Oe) = 0
for all c of given J , M,
(9.2)
which is of course independent of the channel selected as entrance channel in the above argument. Together with the wave equation
H~=
ko~+Eso
)
~
(9.3)
a n d the condition of finite probability IWI2 finite in the interior region,
(9.4)
the b o u n d a r y conditions (9.2) will therefore characterize a set of eigenfunctions W:M, ,, which are not related to a n y particular entrance channel: the corresponding complex wave numbers k], 0n are of course independent of M. The radial factors of the eigenfunction TJM, n are also independent of a n y entrance channel and m a y accordingly be designated b y ~],en, q}Y,en" To the wave numbers k.L on of channel 0 are associated, b y means of eq. (8.7), a set of wave numbers k], cn for a n y other channel c, and a set of complex energy eigenvalues 81, n satisfying energy conservation relations h2
°a], n :
h2
2 M ° kj,~ on'~E~o  2M e k}, cn+ E;o . . . . .
(9.5)
F r o m eqs. (7.2) and (9.2), we see t h a t all the m a t r i x elements U~,°,z,,z a corresponding to a fixed J have the same poles k], on in the physical k0plane. Since the ke are singlevalued functions of ko, we need not always refer to the J channel 0, b u t m a y also say t h a t the poles of the U~,8,,, ' ~o~ are the k], e~. F r o m now on, we shall usually omit the indices J M in keiM' n and the index J in k], en and @:, n . 9.2. PROPERTIES F r o m eqs. (8.12) and (8.9), it is evident that, as in the onechannel case, with each pole ke. = KeniTe ~
(9.6a)
is associated the pole  he* =  Kc, ire,,
(9.6b)
~70
J.
HUMBLET
A N D L. R O S E N F E L D
and we m a y therefore take without restriction Ken ~ 0.
(9.6C)
The eigenfunction belonging to the eigenvalue k0.* will be denoted as ~ ] U ,  . , t and its radial factors as #y, c., ~], c.. On account of the timereversal relations (8.12), the latter are related to the radial factors #],e., #],e. by ~],' e. = ~J.'* c,*"
qbl, c. = q5~,c.,
(9.7)
Writing 8 . in the form e, = E.½ir.,
(9.8)
we have, on account of the energy relations (9.5), h2 E.

'K2
~ '+E
2 M ° ( o.7oM
~o
2~ 2
/'~


/P 2M c
2 ~ +Ezo (%,7c,)  . . .,
(9.9a)
2~ 2
Mo %.70.

M e Kc"Ye" . . . . .
(9.9b)
For a decaying state (F. > 0), the last relation corresponds to the fact that the rate of decay is the same in all channels, only the branching ratios being different for the various channels. It m a y again be shown that Yc. >= 0
when
Kc. :/= O,
(9.10)
which, by (9.6c) and (9.9b), amounts to proving that F , ~ 0.
(9.11)
This results from the actual computation of the width /',, carried out by means of the form (8.10) of Green's theorem, applied to k~1  k~2 = ku,. On account of (9.8) we have f'~
~ee~2
{~:"~e.
~e*''~:*t
(9.12)
In virtue of the conditions (8.2), the quantity between brackets m a y be written, with the notation (3.22) (in which the channel symbol c is used rather than the quantum number l belonging to it), Le(ke,)Le(ke,)*. For this quantity we have the alternative expressions (3.24) and (3.27), which, on account of eqs. (9.6), (9.9b), m a y be put into the form IOe.]2ELc(kc.)Lc(kcn)*] = 2iKc.]~0 m (ke.bc)[ 2 i~_r'~
If? c
I0~(%, ko.) 12dr~t Jg'o~ (b~)J,
(9.13)
THEORY OF NUCLEAR REACTIONS (I)
2Me/~. N 10e.l~[Lo(k¢.)L¢(k,.) *] = 2 i K c .   i  ~ . on"
571
(9.14)
Substituting the expression (9.13) into eq. (9.12), we obtain F n as the quotient of two essentially positive quantities, which proves eqs. (9.11) and (9.10). As in section 3, however, we prefer, for actual use, the expression for Fn derived from the substitution of (9.14) into eq. (9.12):
11.
~.(#~'<°,,IMe)l@°nlO°.12 <
(9.15)
f ,o I~"l~d°~+ Z Ncnl@e"lO~"l= c
in which, b y the same argument as in section 3, the denominator is an essentially nonnegative quantity. The classification of the resonant states discussed in section 2 for the singlechannel case can be performed b y considering the values of ken which correspond to a given location of ko~ in the physical koplane. A state t h a t is "virtual" in the channel 0 (i.e. K~, < Yo*n) has the same property in all the other channels (i.e. Ken 2 < y2~); the proper resonant states (K2n > y2n), on the other hand, appear as virtual states in the channels c with internal energy E~c higher than the resonance energy E~, because then Ken2< Yen'2 The occurrence of real wave numbers k0n  Ken > 0 m a y be excluded b y the following physical argument: such a solution of the eigenvalue problem would correspond to an actual stationary state of the system with pure outgoing waves in open channels and no incoming wave in any channel; this is in opposition with the condition (9.4). As in section 1, we disregard the very special case corresponding to ken = 0. We m a y then draw the cuts in the k0plane so close to the real axis that they all lie above some line k0  A, where A denotes a positive q u a n t i t y smaller than all the distances Yon of the points k0, from the real axis. Under such conditions, it m a y also be seen that k0n is uniquely determined in terms of ken. We still have to consider the states of real (negative) energy En, corresponding to K0~  0: they have in all channels the property Ken = 0. Likewise, the imaginary part Yen of the channel wave number has the same sign in all channels, and the state in question is thereby unambiguously characterized either as a bound state or as a virtual state of the "deuteron" type.
10. E x p a n s i o n of the Collision Matrix 10.1 ANALYTICALPROPERTIES OF THE COLLISIONMATRIX The analytical behaviour of the matrix elements Uc, c, according to their definition (7.2), depends primarily on that of the radial factors ¢(~), q~¢c ?)' of the interior wave function ~c~. Considering the way in which we have derived ~C~(ko) from TcCl(ko, kc, k c. . . . . ) in section 8, we must expect
572
J. HUMBLET AND L. ROSENFELD
• (cc,l fk0), ~Ccc!'(ko) to have branch points in the k0plane at the thresholds + K e , +Ke,, . . . . We assume t h a t these singularities are the only ones the radial factors have for finite ko, and t h a t t h e y are not simultaneously zeros of these functions. The absence of zeros at the thresholds is at a n y rate in agreement with   t h o u g h of course not a consequence of   the timereversal relations (8.4). Under such assumptions, in the physical koplane, i.e. the k oplane with its cuts, the radial factors are singlevalued and regular analytic functions. It m u s t be observed t h a t the introduction of such a hypothesis of a n a l y t i c i t y is m a d e possible in particular b y the circumstance t h a t the collision m a t r i x is independent of the normalization of T (e). This assumption about the analytical behaviour of the radial factors is not required (and indeed not fulfilled) when we consider, for instance, the normalization of Tce~ to unit incoming flux (cf. eqs. (1.21) and (6.8b)) in order to calculate the crosssections in terms of the collision matrix; or when we temporarily use some other normalization, such as in eq. (7.8). Let us now consider the collision m a t r i x element Ve, e itself. In the denominator, the function 0e, like ~ z (1~, has a pole of order l in ke = 0, and conseq u e n t l y Ve, e has a zero of order l at t h a t threshold. Turning to the Wronskian of the numerator, it seems at first sight t h a t it has, like Ie,, a pole of order l' in k e, = 0. It is readily seen, however, t h a t this is not the case. For this purpose, in analogy with w h a t we did in subsection 1.4, let us define the q u a n t i t y Se, e = V M e
1
(Ue'c~e'e),
+1
(10.1)
~ M e,
which we first consider for c' = c. We have 1 ie) See = ek'+l W(~be(e)' keOe)* ,
(10.2)
where, according to eqs. (3.15), we use the function ie = z o  o e
= 2i
d.1,
(10.3)
which, like J~(keac), is the product of ke z+l b y an integral function of ke2. I t is thus clear t h a t See is finite for ke = 0. For c' va c, on account of eqs. (9.1b), we m a y again write
Sc, c
W~e =
1W " ck~ .c'kV+~
(~(¢¢')' I~'0c') W(q~(cc) , Oe)
=
V ~cc
l
M e ' e"
W (~(~') ' Je')
(10.4)
, e,
and therefore conclude t h a t in a n y case Se, e is finite in k e = 0 and k e, = O, i.e. t h a t Ue'e~c,e behaves according to Ve, e(~e,c
near these thresholds.
G<: 7.~+1~' rve ~e ~
(10.5)
573
THEORY OF NUCLEAR REACTIONS (I}
The general conclusion of the preceding analysis is thus t h a t the matrix element
ic,/kc,)

so,o =
=
VMo,
'
(10.6)
which directly governs all the reaction crosssections, including those for elastic scattering, has no other singularities in the physical k0plane than poles corresponding to the resonant states. It can accordingly be expanded in this plane in the form indicated by the MittagLeffler theorem. It is more advantageous, however, to use an expansion in the "physical" plane of the total
v
Fig. 3. T h e p h y s i c a l oaplane w i t h t y p i c a l c u t s f r o m c h a n n e l t h r e s h o l d s . B = b o u n d s t a t e , V = v i r t u a l s t a t e of d e u t e r o n t y p e .
energy e , along the same lines as in subsection 2.3. As functions of ~, the matrix elements Se, c have branchpoints on the positive real axis at the thresholds ~ = E~ of the various channels. It is convenient to choose as the origin of the energy scale the threshold of channel 0, which amounts to putting Es0 = 0. The physical 8plane is then defined b y drawing cuts from each of the thresholds (including 6~  0) along the real axis towards  oo. The resonant states are represented b y points in the lower halfplane and possibly also on the negative real axis, the latter corresponding either to bound states or to virtual states of the "deuteron" type. The cuts should pass below the boundstate points and above the "deuteron"state points and be so close to the real axis t h a t no resonant state should lie on any of them or between two of them (fig. 3).
574
j. HUM~=~ . N . L . ~OS~SFELD
10.2. EXPANSIONS
OF
THE
COLLISION
MATRIX
W h e n the choice of the i n d e p e n d e n t variable has been decided, the arbitrariness in the convergence e x p o n e n t occurring in the MittagLeffler formula still leaves room, in the m a n y  c h a n n e l t h e o r y as well as in the singlechannel case, for a v a r i e t y of expansions of the same general t y p e for the m a t r i x elements S¢,c, Se'c. I t is reasonable to assume t h a t the a s y m p t o t i c location of the r e s o n a n t states will h a v e the same c h a r a c t e r for the general case of nuclear i n t e r a c t i o n as for a fixed potential, and t h a t the m i n i m u m admissible value of the convergence e x p o n e n t M will be the same as in the l a t t e r case. Here, we shall discuss the expansions of b o t h Sc, c and Se, c with the convergence e x p o n e n t M = 0; as in subsection 1.4 the first expansion will bring out explicitly the threshold behaviour, while the o t h e r has a somewhat simpler form. The residue rc, en of Sc, c at oz = @~ is rc, c, =
V ~~ ' ~ ,
k l ' + l ~w(~o,., . . . io,.) .
c.c,.
'
(10.7)
_ _ ~ w ( + ¢ (c, , oc)] Ltxo
_j 6, = oan
where/'c'~ = Ie, nOc,,,. T a k i n g account of eq. (3.18) and of the conditions
W@c, ., Oc,,) = 0
for all c',
(10.8)
satisfied b y the radial factors q~c,n, #'c'~, we get ~Ctn
W(qSe'n, re'n) = 2ikc'n Oc,~.
(10.9)
I n order to calculate (d/doa) W(~v (c) c , 0o), we s t a r t from the relation (8.10) derived from Green's theorem, in which we t a k e T 1 ~ ~cc) (k), ~ = 7 t(e) (k*,,) = hu_, (the n o t a t i o n k represents the set of channel w a v e  n u m b e r s kc, kc,, • • • belonging to the e n e r g y g ) :

c' 2 M e '
xe'n,
I n v i r t u e of the eqs. (9.7), the W r o n s k i a n on the r i g h t  h a n d side is successively t r a n s f o r m e d into
[ +t~',' Lo,.] w(+o.., +~,,(k)) = ¢o..+t~,,(k) L~:.' when account is t a k e n of eqs. (10.8) and the n o t a t i o n (3.22) is used. T h e last expression m a y f u r t h e r be split as follows:
w(¢o,., ~(2(k)) = +o,.¢¢V(k)[Lo,(k)L~,.]
+°'" W~¢
T H E O R Y O F N U C L E A R R E A C T I O N S (I)
575
the last term vanishing, in virtue of the conditions (9.1b), for all channels except c. We therefore get
f
~
t~2
~c~
 
W @X C(~) 0~) ,¢,
We now take the derivative of both sides with respect to 6~ and let then ~(c) (k) coincide with ku~; it is clear, on account of the conditions (10.8) and the special form of the various terms, that the result of this operation is g ~ ~e'*[ d W¢~(c) Oc)] + ~ ~ ~2 dLc'  J~ ~'*,~ ~,,dco 2M c Oc,, ~~ , c , ~'= e,, ,2Me, c',, ~ t = 0%' whence 1
h2 ¢ c .
1
2Me Oe~ vn '
(10.10)
with vn = f~ ~*,~krJndc°+2 ~2__ ~cZ,, dLe~ c 2Me d@n"
(10.11)
Inserting the results (10.9) and (10.10) into eq. (10.7), we obtain for the residue re, e~ the symmetrical expression ~i2 ~CsCn ~
i
1
qbc. qbe, n
l l' V'McMc, vnkcnkc'n Ocn Oe'n"
(10.12)
We now define by ~bcn
e'~.~o~  s / ~  ~ ~ oc~
(10.1Sa)
the modulus Gc~ and phase ~o~ of one of the typical factors occurring in the expression (10.12). Denoting, moreover, by Ce~ the phase of the complex quantity k~c~, and defining ~c~ by ~c~ = ~c~¢e~,
(lO.13b)
we have the expansion Sc,~
eqo ". Ge, nGen e qo.
(10.14a)
The function Qe, e(d~) is regular in the physical o~plane; it is undefined along the cuts, and in particular at the thresholds. A similar expansion is valid for So'e, namely ~o,~ : Oo,~(~')i $. e'~o'. Go, ~ G ~ e qe . (10.14b)
576
J. HUMBLET AND L. ROSENFELD
Since the collision matrix elements and their poles 6~, are aeindependent quantities, the same property is valid for the residues (10.12)and accordingly also for the quantities Go,, ~e, and ~e,. Corresponding to the expansions (10.14), we have for the collision matrix elements
i ~ ~[~,.I) ~lk~e.[] e °'" 8~,,
'
¢//e,eOe, e = "x,/ke~ [Qe, e(d~)i Ze '~'0',, Gc"*Ge" e '~0] ,, ,of' ~,, " 10.3. P K Y S I C A L
INTERPRETATION
O F TFLE C O L L I S I O N
MATRIX
(10.15a)
(lO.15b)
EXPANSION
Let us return to the expression (9.14) for the total width F,. Defining Ft, by re. =

1 ~2Kcn ]~ 2, . 
ff~ Me
(10.16)
where ft. f
17s.[2dw+ ~ U e .
q)e. 3,
(10.17)
we m a y rewrite the equation (9.14) in the form /'. = ]~ Fe.;
(10.18)
o
clearly, F¢. has to be interpreted as the partial width of the state 7s. in channel c. A penetration/actor Pc. in channel c for the resonant state }/'. is appropriately defined by eikn a
Pe"(ac)  Oc(ae, k.)
2
:
(10.19a)
for an sneutron channel, it reduces exactly to unity, and in general it remains finite for any choice of the channel radius ae. Accordingly, the quantity g~.
F e . _ 1 h2Ke. Pc. fin Me e2Z'c"%[~bcnl2
(10.19b)
is the reduced width of channel c for the state }P.; the quantity go. itself, taken as positive, is the corresponding reduced width amplitude. The partial width Pc., like the total width F., is an aindependent quantity, i.e, it does not change when any of the channel radii a c, ac,,.., is increased from its smallest possible value (of. end of section 3). The penetration factor and reduced width of a given channel, on the other hand, are of course dependent on the choice of the corresponding channel radius, except in the case of an sneutron channel.
THEORY
OF
NUCLEAR
REACTIONS
(I)
577
The expansions (10.15) are easily rewritten in terms of the above quantities. Introducing the notation qn
~
#n
(10.20)

to designate a factor which is channelindependent, we have c o . = q.
= q.
go..
(10.21)
Hence, instead of the expansions (10.15), we get
qle,ebe, c = k~c;+~k~e+JQe,e(@) ._. [ k c , ~ U[ k o ~ Z V _kc, _ kc ei~o,n 1/ ke, ke ~le, e6o, e = Vk~,ke Q e , e ( ~ )  i X qn V e q~'n
t¢C'n Ken
c'n
ei~en,
cn
(10.22a)
i
cn eqo." (10.22b) o~O~n
In these expansions, the factors Fie. m a y be replaced b y P ~ ge~, and the denominators of the resonance terms m a y be given the familiar form
##,, =
EoE¢.+{iF. =
(10.23a)
Ee,Ee,,,+½iF,,,
where Ee
m
2Me k°2,
__
9.
2
Ec, = E,~E~o  2M~e (Kc"Tc")"
(10.23b)
The quantities Kc, and ~'c. are easily expressed in terms of the resonance energy parameter Ec, and the total w i d t h / ' , . The resonant states, as we define them, have a partial width in each channel c, independently of the sign of Een, i.e. even if 6~, appears as virtual in the channel c because Ec, < 0. This corresponds to the fact that the welldefined physical situation represented b y a resonant state (with F , > 0) is b y no means a stationary state of definite positive energy d~ of the system, b u t must be regarded as a superposition of all the states T(d") of positive energy. Conversely, in order to represent a single stationary state, we need the whole set of resonant states. Besides the resonance contributions, all elements of the collision matrix contain in principle background terms Qe, e (°°), Qe, c (8), which, like the resonance part, are independent of the choice of the channel radii (in the sense explained at the end of section 3). It seems difficult to make any general statement about the relative importance of such background terms. As regards their behaviour as functions of the energy, it is clear that, since they depend on
J.
578
HUMBLET
AND L. ROSENFELD
oa through the channel wavenumbers ke oc (o*E~) t, they will exhibit "cusps" at the channel thresholds. In this way, this wellknown feature of the reaction crosssections also finds its natural place in the theory. Our thanks are due to Prof. R. Peierls and to Dr. H. Weidenmtiller for helpful discussions. One of us (J.H.) wishes to express his gratitude to the Copenhagen Institute for Theoretical Physics and the Kellogg Radiation Laboratory at Pasadena, which have offered him excellent opportunities for his work; grants from the University of Liege and the Institut Interuniversitaire des Sciences Nucl6aires are gratefully acknowledged. References I) H. A. Bethe, Rev. Mod. Phys. 9 (1937) 6 9 ; H. A. Bethe and G. Placzek, Phys. Rev. 51 (1937) 450 2) A. M. Lane and R. G. Thomas, Rev. Mod. Phys. 30 (1958) 257 3) P. L. K a p u r a n d R. E. Peierls, Proc. Roy. Soc. A 166 (1938) 277 4) A. F. J. Siegert, Phys. Rev. 56 (1939) 750 5) J. Humblet,~ Thesis, University of Liege (1952); M6m. Soc. Roy. Sc. de Libge (8 °) 12,
No. 4 0952) 6) 7) 8) 9) 10)
ll) 12) 13) 14) 15) 16)
R. G. Newton, J. Math. Phys I (1960) 319 R. E. Peierls, Proc. Roy. Soc. A 253 (1959) 16 K. J. Le Couteur, Proc. Roy. Soc. A 256 (1960) l l 5 L. Rosenfeld, Foundations of the optical model and direct interactions, Ilercegnovi Summer meeting, 1959 (NORDITA, Copenhagen, 1960) Proc. Int. Conf. on Nuclear Structure held in Kingston, Ontario, Canada, August 1960; ed. b y D. A. Bromley and E. W. Vogt (University of Toronto Press, Toronto and North Holland Publishing Company, Amsterdam, 1960) H. M. Nussenzveig, Nuclear Physics I I (1959) 499 G. Beck and IL M. Nussenzveig, Nuovo Cim. 16 (1960) 416 F. J. Shore and V. L. Sailor, Phys Rev. 112 (1958) 191; E. Vogt, Phys. Rev. 112 (1958) 203 E. T. Copson, Theory of Functions of Complex Variables (Oxford University Press, 1935) §6.8 J. llumblet, Comptes rendus Acad. Sc. 231 (1950) 1436 C. Bloch, Nuclear Physics 4 (1957) 503