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Triviality of the S-matrix in the quantum Liouville field theory
Volume 148B, number 1,2,3
PHYSICS LETTERS
22 November 1984
TRIVIALITY OF THE S-MATRIX IN THE QUANTUM LIOUVILLE FIELD THEORY T. YONEYA 1 CERN, Geneva, Switzerland
Received 24 May 1984
Within the framework of the translation non-invariant quantization suggested recently by D'Hoker, Freedman and Jackiw, it is established that the exact S-matrix of the quantum Liouville field theory is trivial. In connection with this resuit, we clarify some geometrical aspects of the theory as interpreted as a model for quantum gravity in two dimensions.
Quantization of the Liouville field theory has attracted much attention. Most of the recent works have mainly been intended for application to string theories [ 1 - 3 ] . However, viewed as an ordinary field theory model in two space-time dimensions, the Liouville system has several peculiar properties [4]. For instance, it seems that a normalizable translation-invariant ground state does not exist. Concerning this problem, D'Hoker, Freedman and Jackiw [5] have made a proposal that the quantum theory be defined on background fields which are translation non-invariant * l They showed that there exists a well-defined perturbation theory, without infrared divergence, around such a background field, and that the tree level S-matrix is trivial. The question whether the S-matrix is trivial in the full-fledged quantum theory is important for understanding the quantum-theoretical significance of the geometrical nature of the Liouville system. This question, however, remained unanswered in their work. In the present note we shall re-examine this question, and establish that the exact S-matrix is indeed trivial if a suitable revision is made in the formulation. As is well known, the Liouville equation 17¢ + (m2/fl) e#~ = 0
(1)
determines a two-dimensional space-time with constant (negative) curvature, in the conformal gauge in which the metric tensor takes the form guy = r/uu × exp(l~0), with r/uv being the flat metric ,2. The conformal invariance of (1) merely reflects the freedom o f coordinate transformation remaining in the conformal gauge. The classical hamiltonian structure corresponding to this geometrical property has been elucidated by Teitelboim [7]. If we decompose the general metric tensor in the form
guy
= e~¢[071) 2 -- (r/l) 2 71
(2)
the generator which describes the deformation of the Cauchy curve at "time" x o to the curve at "time" x o + 8x ° is given by H 6 x ° : H = f d x 1 [n±(x) 2 ± ( x ) + nl(x) 21(x)1 ,
(3)
x 2 + ¢,2) _ (2//~)¢" + (m2//32) e ~¢ 2 l = 5(rt
(4a)
2 1 = ~'¢' -- (2//3)rt',
(4b)
where ~r(x) is the conjugate momentum for ~ x ) . The 2 ± and 2 1 satisfy the following Poisson bracket algebra: [ 2±(x), 2±Qy)] PB =
1 On leave from the Institute of Physics, University of Tokyo, Komaba 3-8-1, Tokyo 153, Japan. ,1 This reminds us of an earlier proposal of Fubini [6] for general conformally invariant field theories.
]1],
t_
ax~(X'Y)[2 1 ( x )
+ 21(Y)1 ' (5a)
,2 We use the Pauli metric, l:] ~ -~)~ ~-ggtZV~ w 111
Volume 148B, number 1,2,3
PHYSICS LETTERS
[ ~ l ( X ) , 9{±(_v)]pB = axS(X,y)[~±(x) + ~±Cv)]
- Za3x6(X, y) ,
(hb)
[~l(X), ~a(.V)]pB = axa(x,y)[ %(x) + ~'10,)], (5c) with a non-vanishing central charge Z = 4/~ 2. The algebra (5) expresses the fact that the evolution from a given initial curve, embedded in a two-dimensional space-time, to a given final curve is independent of the paths of deformation. By choosing r/± = 1 and 7/1 = 0, we find (3) to be the LiouviUe hamiltonian. A consequence of this and the Hamilton equation of motion is that cjg± and ~ 1 satisfy 9~± = ~ ] and 9g i = ~ 1 and hence are free massless fields. This also reflects that ~ ± and ~ 1 coincide with the diagonal and offdiagonal components, respectively, of the improved energy momentum tensor o f the Liouville system. Because of these properties, the Liouville model is also attractive as a possible model for quantum gravity in two dimensions [7,8]. Now the observation that ~ t and ~ 1 satisfy the free field equation will immediately lead to the triviality o f the classical (or tree level) S-matrix [5] provided that the asymptotic field is massless and couples with 9g± or ~ 1 . It is thus reasonable to expect that the preservation of (5) at the quantum level will result in the triviality of the exact S-matrix. It is in fact akeady known that, if one naively defines a quantum theory by performing the normal ordering [2,4] with respect to a free translation-invariant Fock vacuum, (5) can be preserved. Unfortunately, however, this does not solve our problem, since (i) as mentioned above it is highly plausible that there exists no normalizable and translation-invariant state in an exact theory, and (ii) it is not known how to define the asymptotic states without knowing the nature of the ground state. In the naive translation-invariant quantization, any perturbative attempts toward checking the properties o f would-be asymptotic states are marred by the infrared divergences in the infinite volume limit. Therefore, let us follow the suggestion made in ref. [5] and examine ,3 whether it is possible to preserve ,3 Another possible approach might be to try to define a finite-volume counterpart of the S-matrix and finally to take the limit of infinite volume. We do not know how to carry out this. 112
22 November 1984
the geometrical structure expressed in (5) within the new scheme. This task is not trivial because there is, at least apparently, no canonical transformation which could connect both schemes. The quantization starts with the splitting ~ = ~s + ~b, where ~os is a classical static solution of (1). In general, ~s is left invariant under the SO(2.1) subalgebra of the conformal algebra. Hence Poincar6 symmetry in ordinary quantization is now replaced by the SO(2.1) symmetry. Following ref. [5], we choose the simplest solution Cs(X 1) = --/3-1 In ~m2r 2 ,
(6)
where r = x 1 - x 1, and x 1 is an arbitrary integration constant associated with the translation invariance of (1). The key feature of this quantization is that there is no normalizable zero mode, and the complete set of the linearized oscillation modes about (6) are given by e +-it°t tbto(r), with ¢)¢o(r) = (1601r)l/ 2J3/2(t601r), which satisfy a~
f d60 ~kto(r) ~to(r ,) = 6(r - r') ,
0
f dr t~to(r)~bto,(r ) = 6(6o - 60').
(7)
0 Thus, unlike the quantization around solitons, Poincar~ invariance is never restored. This should be expected because the Poincar6 group is not a symmetry o f space-times with constant (non-zero) curvature. Eqs. (7) further show that the representation of the canonical commutation relation (CCR) should be built on a half (infinite) space, e.g. r / > 0. Since ~ko~(0) = ~k~(0) = 0, there is no communication to the other half space. We therefore postulate the following expansion as a representation of CCR:
~(,)= j~7 d 6 0 , ~,,,~*,,:)~,,~(,), o
{fir) = - i j d 6 0 (6012)112(ato- a:)~co(r), 0
a
a t 1 = ~a(60 -- 603
(8)
and def'me the normal ordering with respect to the creation and annihilation operators a ~ and a¢o which
Volume 148B, number 1,2,3
PHYSICS LETTERS
1 [~2 + ~ ' 2 diagonalize the free hamiltonian H (0) = f~o "~ + 2~2/r 2 ] dr. This prescription removes all ultraviolet divergences in the loop expansion. Also, there is no infrared divergence because ~w(r) ~ O(co 2) as co ~ 0. We now come to the main question: What are the deformation generators ~ u satisfying (5)? We may first try to see whether the simple substitution ~0~ % + ~ in the normal ordered expression of ref. [4] meets our requirement. The answer is no. Thus there is no canonical transformation, even formally, which connects both quantization schemes. Fortunately, however, a slight modification can be shown to work. Let us first define
A±(v) -- it(r)+- ~'(r) .
(9)
Formulas (9) satisfy the commutation relations t
_
t
r )[Le(v)
+
[A~+)(r), ~(-)(r')] _1
- ~
oo
.
fdco[-lqj~o(r)~(r 0
~
)+- co
-1
~
~(r)~
t
(r)]. (llb)
It is not difficult to see that (11) have the following short-distance behaviours: [A (+)(r), A (-)(r')] = (h/~r)[-(r - r'g ie) - 2 + 2(r + r ' ) - 2 + S(r-r')] ,
(12a)
[a(.+)(r), 6(-)(r')] = (h/21r)[-7-(r - r'V- ie) -1 + (r + r') -1 + Y,(r-r')] ,
2 (d/dr + 1/r) A~(r), L+(r) ~ :-~A+(r)2: -Y-~
[L±(r),L+(r )] = +-2inbrg(r
22 November 1984
(lZb)
where the remainder parts S(r-r') and Y,(r.r') are real and satisfy S(r'r') = S(r'.r) = O[(r - r') z, (r - r') 2 X log(r - r ' ) ] , Y,(r-r') = O [r - r', ( r - r') l o g ( r - r')]. Then, using the formulas (X - y ~ ie) - n - (x - y + ie) - n
Le(r')]
= +27ri(--) n - I [(n - 1)!] - 1 8 ( n - 1 ) ( x - y ) ,
+ i~ ~ ( O [ O r + 1/r)(O/br'+ 1/r')arf(r - r') + _~([A~+)(r), A(.-)(r,)l 2 _ [A (+)(r,), A(.-)(r)] 2 ) , (a0a) [L+(r), L_(r')] = 0 ,
(10b)
we find that the last terms of the rhs of (10a) and (10c) take the form T-ih2@6(r - r')/6n + 4 i h 2 a r 6 ( r - r')/ rt(r + r') 2 and -T-i/~2(~ r + r - 1)8(r - r'): ep¢(r'):/4rt, respectively. Putting all these results together, we establish that ~(r)
[L+_(r), :eP ~(r') :] = - i h 6 (r - r'): ep~(r')pA +_(r'):
= :~1 [/r(r) 2 + ~,(r)21 : _ (2//3)q~"(r) - (2/~)q~'(r)
+ (2/(32r2)(m 2/#2): ep~(r): + o/r 2 ,
+ ih ~ - ( b / a r + 1/r)6(r - r'):ep~(r'):
~(r)
+ ~p2([A(+)(r) ' tb(-)(r,)]2 (10c)
_ [A(.-)(r), ~b(+)(r')] 2):eP~(r'): ,
- :/r(r)~'(r): -(2//3)[/r'(r) + r-lit(r)]
(13a) (13b)
satisfy the required commutation relations corresponding to (5), provided that
where the superscripts(+) and (-) denote the annihilation and creation part, respectively.In terms of ~to(r), the commutators on the rhs of(10a) and (I0c) are given by
p/[3 - p2~/8rt = 1,
(14a)
z = 4//32 + h / 1 2 n ,
(14b)
[A(+)(r),A(-)(r')]
o = -2//32 + ti/4,r.
(14c)
1
o*
+.
t
t
-T-"
t
t
= ~n f dco[-l~,o(r)q,,o(r ) xq,,o (r) ~ co(r ) 0 ' ' ' + co~¢o(r)~oo(r ' ) + co- 1 qJw(r)~co(r )] ,
(11a)
It should be noted that both the condition (14a) and the quantum contribu$ion to the central charge coincide with those obtained in the translation-invariant quantization in a finite (periodic) box [2], but not with the results of ref. [4]. The undetermined constant g2 in (13a), corresponding to the arbitrariness of 113
Volume 148B, number 1,2,3
PHYSICS LETTERS
constant field translation, is fixed by demanding that the vacuum expectation value of ~ be zero. Then,/a2/ m 2 = 1 - g/325/4rt with ~ "~ O(1) because of the existence of the ~-dependent linear term in the hamfltonian. In the classical limit, formulas (13) are identical with (4). For non-zero h, however, the hamiltonian constructed from (13a) does not coincide with the classical one or with the one given in the literature [2 - 5 ] . The expansion in h of (13a) gives an infinite number of h-dependent terms in the interaction hamiltonian. These terms, which are absent in ref. [5], should be interpreted as the (finite) counter terms which must exist in order to preserve the full geometrical significance of the classical Liouviile theory at the quantum level. As shown before, establishing the existence of the deformation generators ~ satisfying the algebra (5) implies that they obey the free (massless) field equations in the conformal gauge. Hence, defining ^
~(r)=(d/dr - 1/r)~Cu(r)
( p = l , 1),
we now have (l-q + 2 / r 2 ) ~ ( r ,
t) = 0 .
(15) ^
Furthermore, it is easy to see that the operators ~ connect the vacuum to one-particle states, and hence can serve as interpolating fields [5]. Consequences of this and the LSZ reduction formula are that, to all orders in h, (i) there is no quantum correction to the eigenvalue of the SO(2.1) Casimir operator r21~ [5] for the asymptotic one-particle states, and (ii) the Smatrix is an identity. As a check, we compute the self-energy correction to the first non-trivial order in h. The one-loop graph and the contributions coming from the h-dependent counter terms are shown in fig. 1. We fred that, on the energy shell, the sum of all the contributions exactly cancels: (a) = --4i/32w/127r; (b) = ~i/32w(1 + 8)/61r; (c) = -/~/326o(1 + 26)/12~r. This explicitly demonstrates the crucial role of the ~-dependent terms in the hamiltonian, without which the Casimir eigenvalue is
h to
~
to (a)
to
h (b)
to
tO
I
to
(c)
Fig. 1. The Feynman diagrams contributing to the self-energy to order h.
114
22 November 1984
subject to quantum correction ,4 We note that although the language of the manifestly SO(2.1)invariant formulation [5] has not been used, the validity of the algebra (5) assures us that the SO(2.1) symmetry of the space-time with constant curvature can be exactly preserved since the central charge vanishes for the SO(2,1) subalgebra. Let us now return again to the geometrical aspect of the theory and discuss about the general coordinate transformation in the theory as interpreted as a model for quantum gravity in two dimensions. A characteristic feature of the hamiltonian dynamics for systems with general coordinate invariance is the existence of the constraints associated with the arbitrariness of the coordinate choice, i.e. arbitrariness of r f . In quantum theory, the constraints are usually interpreted as the conditions on permissible state vectors, in analogy with the Klein-Gordon equation for the case of a relativistic point particle. The resulting system of equations replaces the Schr~dinger equation of ordinary quantum mechanics, and is called the Wheeler-DeWitt equation [9]. In our case, however, the presence of the non-zero central charge in (5) prevents us from doing this. Although, from the point of view of the integrability of the deformation of the Cauchy curves in two dimensions, the central charge is allowed to exist [7], the resulting non-invariance of the action under general coordinate transformation would, in general, imply that the physical amplitudes change when the space-time parametrization is changed. There is a natural way of remedying this situation. Given an auxiliary free scalar field ×(r) and its conjugate momentum P(r), the operators
1 2 + ×,2): + (2/~)×", ~7~ = - : ~(P
~
=
:Px': -(2/~)P',
(16a) (16b)
satisfy the algebra (5), which is the same as for (13) but the opposite sign for the central charge because of the "wrong" sign in (16a). Here, we assumed the following CCR representation, which diagonalizes the "hamiltonian" f 9 ~ dr:
,4 In ref. [5], it is asserted that the Casimir eigenvalue is unchanged by the quantum correction without the h-dependent terms. There seems no reason for expecting this, and the explicit calculation indeed shows the contrary: (a) ~ 0. The SO(2.1) symmetry demands only the vanishing of the imaginary part of the one-loop graph.
PHYSICS LETTERS
Volume 148B, number 1,2,3 x(r) = / 0
d~,~(2/Tr)l/2cos~°r(bto+b~), (2~°) ~t~
P(r) = - i f d ~ ( e o / 2 ) l / 2 ( 2 f i r ) 1/2 cos 0
6or(b w - b ~ ) ,
[b o, b ~ , ] = h 6 ( w - w ' ) .
(17)
^
The sum ~gu =- ~gu~ + 9~x therefore satisfies the algebra (5) without the central charge. That this is a natural (and almost compulsory) way o f restoring the general coordinate invariance can be seen b y taking the classical limit. The action A = f d2x [~rr + ~(P - ~7u ~gv] is then rewritten as
A-f~(~g _
1
~V
auxax A+R2)d2x ~
' ~
(18)
(A = m 2) w i t h ~ = X - ~0. Formula (18) is just the local version of the non-local invariant action, written down b y Polyakov [1 ] , which is reduced to the Liouville action in the conformal gauge. The hamiltonJan formulation o f the non-local action necessitates the introduction o f the auxiliary field X", which is dynamical. Formula (18) then leads to the constraints ~g± = ~ 1 = 0 in the standard way . s . We emphasize that the "wrong" sign in (16a) does not do any immediate harm, since 9~± is a constraint rather than the energy density. It is amusing to note that the classical general solution o f the Liouville equation in terms o f the free field is, if expressed in the canonical language, nothing but the solution o f these constraints. In this sense, the Wheeler-DeWitt equation itself is a natural quantum generalization o f that classical solution. It is interesting to study the solution o f the Wheeler-DeWitt equation in the light o f this observation. However, it is b e y o n d the scope o f the present note and is left to a future publication. Finally, we add a few remarks related to our generally covariant interpretation: (i) The standard counting rule o f the dynamical degrees o f freedom for the constrained system tells us that there is no true dynamics in the system. It is therefore natural that the S-matrix o f the Liouville system is trivial. (ii) The energy o f a generally covariant system must be expressed b y a surface term, following the standard method [11 ]. It is . s The action (18) has previously appeared in the literature [ 10] where, however, it was only used as a formal device for deriving the improved energy momentum tensor.
22 November 1984
easily seen that the only possible (t'mite) value o f the energy o f our system is then zero, at least classically. A closely related (but not identical) phenomenon will be discussed, in the context o f a different model, in another paper [12]. (iii) In terms o f ~ and ~ = rr + P , the constraints in the classical theory take the form ~ ± = 2 - 1 ~ 2 - ~ P + 2 ~ - l ' x " - 2 - 1 " X '2 - ~0"X' + A exp(~/~0) and ~gl = X r + ~0, r - ztJ- 7r . From these expressions, we observe that in a recent work [13] b y Banks and Susskind only the special case where P = ~ ' = 0 was treated. F r o m our point o f view, this is not, however, a gauge choice. This explains why these authors had only a collapsed solution for A > 0 (negative scalar curvature). It should also be remarked that with the above restriction we have now [ ~g±(x), ~g±(v)] = 0, which may be interpreted as an indication that the s p a c e - t i m e s swept out b y the deformation generators with this restriction are in fact something different [14] from the ordinary (pseudo-)riemannian space-times. I would like to thank A. Neveu for his comments on the manuscript. [1] A. Polyakov, Phys. Lett. 103B (1981) 207. [2] T. Curtright and C. Thorn, Phys. Rev. Lett. 48 (1982) 1309; E. Braaten, T. Curtright and C. Thorn, Phys. Lett. l18B (1982) 115. [3] J.L. Gervais and A. Neveu, Nucl. Phys. B209 (1982) 125; B224 (1983) 329; CERN preprints TH 3731 andTH 3732. [4] E. D'Hoker and R. Jackiw, Phys. Rev. D26 (1982) 3517. [5] E. D'Hoker, D. Freedman and R. Jackiw, Phys. Rev. D28 (1984) 2583; E. D'Hoker and R. Jackiw, Phys. Rev. Lett. 50 (1983) 1719. [6] S. Fubini, Nuovo Cimento 34A (1976) 521. [7] C. Teitelboim, Phys. Lett. 126B (1983) 41. [8] R. Jackiw, in: Ahead of his time;Bryce S. DeWitt - a collection of essays in honor of his 60th birthday, ed. S. Christensen (Adam Hilger, Bristol, 1983). [9] See, for example, B.S. DeWitt, Phys. Rev. 160 (1967) 1113. [10] R. Marnelius, Nucl. Phys. B211 (1983) 14. [11] T. Regge and C. Teitelboim, Ann. Phys. (NY) 88 (1974) 286. [ 12] T. Yoneya, Higher derivative quantum gravity in two dimensions, CERN preprint TH 3867, to be published in Phys. Lett. B. [13] T. Banks and L. Susskind, Stanford Univ. preprint ITP743 (1983). [ 14] C. Teitelboim, in: General relativity and gravitation: one hundred years after the birth of Albert Einstein, ed. A. Held (Plenum, New York, 1980). 115
PHYSICS LETTERS
22 November 1984
TRIVIALITY OF THE S-MATRIX IN THE QUANTUM LIOUVILLE FIELD THEORY T. YONEYA 1 CERN, Geneva, Switzerland
Received 24 May 1984
Within the framework of the translation non-invariant quantization suggested recently by D'Hoker, Freedman and Jackiw, it is established that the exact S-matrix of the quantum Liouville field theory is trivial. In connection with this resuit, we clarify some geometrical aspects of the theory as interpreted as a model for quantum gravity in two dimensions.
Quantization of the Liouville field theory has attracted much attention. Most of the recent works have mainly been intended for application to string theories [ 1 - 3 ] . However, viewed as an ordinary field theory model in two space-time dimensions, the Liouville system has several peculiar properties [4]. For instance, it seems that a normalizable translation-invariant ground state does not exist. Concerning this problem, D'Hoker, Freedman and Jackiw [5] have made a proposal that the quantum theory be defined on background fields which are translation non-invariant * l They showed that there exists a well-defined perturbation theory, without infrared divergence, around such a background field, and that the tree level S-matrix is trivial. The question whether the S-matrix is trivial in the full-fledged quantum theory is important for understanding the quantum-theoretical significance of the geometrical nature of the Liouville system. This question, however, remained unanswered in their work. In the present note we shall re-examine this question, and establish that the exact S-matrix is indeed trivial if a suitable revision is made in the formulation. As is well known, the Liouville equation 17¢ + (m2/fl) e#~ = 0
(1)
determines a two-dimensional space-time with constant (negative) curvature, in the conformal gauge in which the metric tensor takes the form guy = r/uu × exp(l~0), with r/uv being the flat metric ,2. The conformal invariance of (1) merely reflects the freedom o f coordinate transformation remaining in the conformal gauge. The classical hamiltonian structure corresponding to this geometrical property has been elucidated by Teitelboim [7]. If we decompose the general metric tensor in the form
guy
= e~¢[071) 2 -- (r/l) 2 71
(2)
the generator which describes the deformation of the Cauchy curve at "time" x o to the curve at "time" x o + 8x ° is given by H 6 x ° : H = f d x 1 [n±(x) 2 ± ( x ) + nl(x) 21(x)1 ,
(3)
x 2 + ¢,2) _ (2//~)¢" + (m2//32) e ~¢ 2 l = 5(rt
(4a)
2 1 = ~'¢' -- (2//3)rt',
(4b)
where ~r(x) is the conjugate momentum for ~ x ) . The 2 ± and 2 1 satisfy the following Poisson bracket algebra: [ 2±(x), 2±Qy)] PB =
1 On leave from the Institute of Physics, University of Tokyo, Komaba 3-8-1, Tokyo 153, Japan. ,1 This reminds us of an earlier proposal of Fubini [6] for general conformally invariant field theories.
]1],
t_
ax~(X'Y)[2 1 ( x )
+ 21(Y)1 ' (5a)
,2 We use the Pauli metric, l:] ~ -~)~ ~-ggtZV~ w 111
Volume 148B, number 1,2,3
PHYSICS LETTERS
[ ~ l ( X ) , 9{±(_v)]pB = axS(X,y)[~±(x) + ~±Cv)]
- Za3x6(X, y) ,
(hb)
[~l(X), ~a(.V)]pB = axa(x,y)[ %(x) + ~'10,)], (5c) with a non-vanishing central charge Z = 4/~ 2. The algebra (5) expresses the fact that the evolution from a given initial curve, embedded in a two-dimensional space-time, to a given final curve is independent of the paths of deformation. By choosing r/± = 1 and 7/1 = 0, we find (3) to be the LiouviUe hamiltonian. A consequence of this and the Hamilton equation of motion is that cjg± and ~ 1 satisfy 9~± = ~ ] and 9g i = ~ 1 and hence are free massless fields. This also reflects that ~ ± and ~ 1 coincide with the diagonal and offdiagonal components, respectively, of the improved energy momentum tensor o f the Liouville system. Because of these properties, the Liouville model is also attractive as a possible model for quantum gravity in two dimensions [7,8]. Now the observation that ~ t and ~ 1 satisfy the free field equation will immediately lead to the triviality o f the classical (or tree level) S-matrix [5] provided that the asymptotic field is massless and couples with 9g± or ~ 1 . It is thus reasonable to expect that the preservation of (5) at the quantum level will result in the triviality of the exact S-matrix. It is in fact akeady known that, if one naively defines a quantum theory by performing the normal ordering [2,4] with respect to a free translation-invariant Fock vacuum, (5) can be preserved. Unfortunately, however, this does not solve our problem, since (i) as mentioned above it is highly plausible that there exists no normalizable and translation-invariant state in an exact theory, and (ii) it is not known how to define the asymptotic states without knowing the nature of the ground state. In the naive translation-invariant quantization, any perturbative attempts toward checking the properties o f would-be asymptotic states are marred by the infrared divergences in the infinite volume limit. Therefore, let us follow the suggestion made in ref. [5] and examine ,3 whether it is possible to preserve ,3 Another possible approach might be to try to define a finite-volume counterpart of the S-matrix and finally to take the limit of infinite volume. We do not know how to carry out this. 112
22 November 1984
the geometrical structure expressed in (5) within the new scheme. This task is not trivial because there is, at least apparently, no canonical transformation which could connect both schemes. The quantization starts with the splitting ~ = ~s + ~b, where ~os is a classical static solution of (1). In general, ~s is left invariant under the SO(2.1) subalgebra of the conformal algebra. Hence Poincar6 symmetry in ordinary quantization is now replaced by the SO(2.1) symmetry. Following ref. [5], we choose the simplest solution Cs(X 1) = --/3-1 In ~m2r 2 ,
(6)
where r = x 1 - x 1, and x 1 is an arbitrary integration constant associated with the translation invariance of (1). The key feature of this quantization is that there is no normalizable zero mode, and the complete set of the linearized oscillation modes about (6) are given by e +-it°t tbto(r), with ¢)¢o(r) = (1601r)l/ 2J3/2(t601r), which satisfy a~
f d60 ~kto(r) ~to(r ,) = 6(r - r') ,
0
f dr t~to(r)~bto,(r ) = 6(6o - 60').
(7)
0 Thus, unlike the quantization around solitons, Poincar~ invariance is never restored. This should be expected because the Poincar6 group is not a symmetry o f space-times with constant (non-zero) curvature. Eqs. (7) further show that the representation of the canonical commutation relation (CCR) should be built on a half (infinite) space, e.g. r / > 0. Since ~ko~(0) = ~k~(0) = 0, there is no communication to the other half space. We therefore postulate the following expansion as a representation of CCR:
~(,)= j~7 d 6 0 , ~,,,~*,,:)~,,~(,), o
{fir) = - i j d 6 0 (6012)112(ato- a:)~co(r), 0
a
a t 1 = ~a(60 -- 603
(8)
and def'me the normal ordering with respect to the creation and annihilation operators a ~ and a¢o which
Volume 148B, number 1,2,3
PHYSICS LETTERS
1 [~2 + ~ ' 2 diagonalize the free hamiltonian H (0) = f~o "~ + 2~2/r 2 ] dr. This prescription removes all ultraviolet divergences in the loop expansion. Also, there is no infrared divergence because ~w(r) ~ O(co 2) as co ~ 0. We now come to the main question: What are the deformation generators ~ u satisfying (5)? We may first try to see whether the simple substitution ~0~ % + ~ in the normal ordered expression of ref. [4] meets our requirement. The answer is no. Thus there is no canonical transformation, even formally, which connects both quantization schemes. Fortunately, however, a slight modification can be shown to work. Let us first define
A±(v) -- it(r)+- ~'(r) .
(9)
Formulas (9) satisfy the commutation relations t
_
t
r )[Le(v)
+
[A~+)(r), ~(-)(r')] _1
- ~
oo
.
fdco[-lqj~o(r)~(r 0
~
)+- co
-1
~
~(r)~
t
(r)]. (llb)
It is not difficult to see that (11) have the following short-distance behaviours: [A (+)(r), A (-)(r')] = (h/~r)[-(r - r'g ie) - 2 + 2(r + r ' ) - 2 + S(r-r')] ,
(12a)
[a(.+)(r), 6(-)(r')] = (h/21r)[-7-(r - r'V- ie) -1 + (r + r') -1 + Y,(r-r')] ,
2 (d/dr + 1/r) A~(r), L+(r) ~ :-~A+(r)2: -Y-~
[L±(r),L+(r )] = +-2inbrg(r
22 November 1984
(lZb)
where the remainder parts S(r-r') and Y,(r.r') are real and satisfy S(r'r') = S(r'.r) = O[(r - r') z, (r - r') 2 X log(r - r ' ) ] , Y,(r-r') = O [r - r', ( r - r') l o g ( r - r')]. Then, using the formulas (X - y ~ ie) - n - (x - y + ie) - n
Le(r')]
= +27ri(--) n - I [(n - 1)!] - 1 8 ( n - 1 ) ( x - y ) ,
+ i~ ~ ( O [ O r + 1/r)(O/br'+ 1/r')arf(r - r') + _~([A~+)(r), A(.-)(r,)l 2 _ [A (+)(r,), A(.-)(r)] 2 ) , (a0a) [L+(r), L_(r')] = 0 ,
(10b)
we find that the last terms of the rhs of (10a) and (10c) take the form T-ih2@6(r - r')/6n + 4 i h 2 a r 6 ( r - r')/ rt(r + r') 2 and -T-i/~2(~ r + r - 1)8(r - r'): ep¢(r'):/4rt, respectively. Putting all these results together, we establish that ~(r)
[L+_(r), :eP ~(r') :] = - i h 6 (r - r'): ep~(r')pA +_(r'):
= :~1 [/r(r) 2 + ~,(r)21 : _ (2//3)q~"(r) - (2/~)q~'(r)
+ (2/(32r2)(m 2/#2): ep~(r): + o/r 2 ,
+ ih ~ - ( b / a r + 1/r)6(r - r'):ep~(r'):
~(r)
+ ~p2([A(+)(r) ' tb(-)(r,)]2 (10c)
_ [A(.-)(r), ~b(+)(r')] 2):eP~(r'): ,
- :/r(r)~'(r): -(2//3)[/r'(r) + r-lit(r)]
(13a) (13b)
satisfy the required commutation relations corresponding to (5), provided that
where the superscripts(+) and (-) denote the annihilation and creation part, respectively.In terms of ~to(r), the commutators on the rhs of(10a) and (I0c) are given by
p/[3 - p2~/8rt = 1,
(14a)
z = 4//32 + h / 1 2 n ,
(14b)
[A(+)(r),A(-)(r')]
o = -2//32 + ti/4,r.
(14c)
1
o*
+.
t
t
-T-"
t
t
= ~n f dco[-l~,o(r)q,,o(r ) xq,,o (r) ~ co(r ) 0 ' ' ' + co~¢o(r)~oo(r ' ) + co- 1 qJw(r)~co(r )] ,
(11a)
It should be noted that both the condition (14a) and the quantum contribu$ion to the central charge coincide with those obtained in the translation-invariant quantization in a finite (periodic) box [2], but not with the results of ref. [4]. The undetermined constant g2 in (13a), corresponding to the arbitrariness of 113
Volume 148B, number 1,2,3
PHYSICS LETTERS
constant field translation, is fixed by demanding that the vacuum expectation value of ~ be zero. Then,/a2/ m 2 = 1 - g/325/4rt with ~ "~ O(1) because of the existence of the ~-dependent linear term in the hamfltonian. In the classical limit, formulas (13) are identical with (4). For non-zero h, however, the hamiltonian constructed from (13a) does not coincide with the classical one or with the one given in the literature [2 - 5 ] . The expansion in h of (13a) gives an infinite number of h-dependent terms in the interaction hamiltonian. These terms, which are absent in ref. [5], should be interpreted as the (finite) counter terms which must exist in order to preserve the full geometrical significance of the classical Liouviile theory at the quantum level. As shown before, establishing the existence of the deformation generators ~ satisfying the algebra (5) implies that they obey the free (massless) field equations in the conformal gauge. Hence, defining ^
~(r)=(d/dr - 1/r)~Cu(r)
( p = l , 1),
we now have (l-q + 2 / r 2 ) ~ ( r ,
t) = 0 .
(15) ^
Furthermore, it is easy to see that the operators ~ connect the vacuum to one-particle states, and hence can serve as interpolating fields [5]. Consequences of this and the LSZ reduction formula are that, to all orders in h, (i) there is no quantum correction to the eigenvalue of the SO(2.1) Casimir operator r21~ [5] for the asymptotic one-particle states, and (ii) the Smatrix is an identity. As a check, we compute the self-energy correction to the first non-trivial order in h. The one-loop graph and the contributions coming from the h-dependent counter terms are shown in fig. 1. We fred that, on the energy shell, the sum of all the contributions exactly cancels: (a) = --4i/32w/127r; (b) = ~i/32w(1 + 8)/61r; (c) = -/~/326o(1 + 26)/12~r. This explicitly demonstrates the crucial role of the ~-dependent terms in the hamiltonian, without which the Casimir eigenvalue is
h to
~
to (a)
to
h (b)
to
tO
I
to
(c)
Fig. 1. The Feynman diagrams contributing to the self-energy to order h.
114
22 November 1984
subject to quantum correction ,4 We note that although the language of the manifestly SO(2.1)invariant formulation [5] has not been used, the validity of the algebra (5) assures us that the SO(2.1) symmetry of the space-time with constant curvature can be exactly preserved since the central charge vanishes for the SO(2,1) subalgebra. Let us now return again to the geometrical aspect of the theory and discuss about the general coordinate transformation in the theory as interpreted as a model for quantum gravity in two dimensions. A characteristic feature of the hamiltonian dynamics for systems with general coordinate invariance is the existence of the constraints associated with the arbitrariness of the coordinate choice, i.e. arbitrariness of r f . In quantum theory, the constraints are usually interpreted as the conditions on permissible state vectors, in analogy with the Klein-Gordon equation for the case of a relativistic point particle. The resulting system of equations replaces the Schr~dinger equation of ordinary quantum mechanics, and is called the Wheeler-DeWitt equation [9]. In our case, however, the presence of the non-zero central charge in (5) prevents us from doing this. Although, from the point of view of the integrability of the deformation of the Cauchy curves in two dimensions, the central charge is allowed to exist [7], the resulting non-invariance of the action under general coordinate transformation would, in general, imply that the physical amplitudes change when the space-time parametrization is changed. There is a natural way of remedying this situation. Given an auxiliary free scalar field ×(r) and its conjugate momentum P(r), the operators
1 2 + ×,2): + (2/~)×", ~7~ = - : ~(P
~
=
:Px': -(2/~)P',
(16a) (16b)
satisfy the algebra (5), which is the same as for (13) but the opposite sign for the central charge because of the "wrong" sign in (16a). Here, we assumed the following CCR representation, which diagonalizes the "hamiltonian" f 9 ~ dr:
,4 In ref. [5], it is asserted that the Casimir eigenvalue is unchanged by the quantum correction without the h-dependent terms. There seems no reason for expecting this, and the explicit calculation indeed shows the contrary: (a) ~ 0. The SO(2.1) symmetry demands only the vanishing of the imaginary part of the one-loop graph.
PHYSICS LETTERS
Volume 148B, number 1,2,3 x(r) = / 0
d~,~(2/Tr)l/2cos~°r(bto+b~), (2~°) ~t~
P(r) = - i f d ~ ( e o / 2 ) l / 2 ( 2 f i r ) 1/2 cos 0
6or(b w - b ~ ) ,
[b o, b ~ , ] = h 6 ( w - w ' ) .
(17)
^
The sum ~gu =- ~gu~ + 9~x therefore satisfies the algebra (5) without the central charge. That this is a natural (and almost compulsory) way o f restoring the general coordinate invariance can be seen b y taking the classical limit. The action A = f d2x [~rr + ~(P - ~7u ~gv] is then rewritten as
A-f~(~g _
1
~V
auxax A+R2)d2x ~
' ~
(18)
(A = m 2) w i t h ~ = X - ~0. Formula (18) is just the local version of the non-local invariant action, written down b y Polyakov [1 ] , which is reduced to the Liouville action in the conformal gauge. The hamiltonJan formulation o f the non-local action necessitates the introduction o f the auxiliary field X", which is dynamical. Formula (18) then leads to the constraints ~g± = ~ 1 = 0 in the standard way . s . We emphasize that the "wrong" sign in (16a) does not do any immediate harm, since 9~± is a constraint rather than the energy density. It is amusing to note that the classical general solution o f the Liouville equation in terms o f the free field is, if expressed in the canonical language, nothing but the solution o f these constraints. In this sense, the Wheeler-DeWitt equation itself is a natural quantum generalization o f that classical solution. It is interesting to study the solution o f the Wheeler-DeWitt equation in the light o f this observation. However, it is b e y o n d the scope o f the present note and is left to a future publication. Finally, we add a few remarks related to our generally covariant interpretation: (i) The standard counting rule o f the dynamical degrees o f freedom for the constrained system tells us that there is no true dynamics in the system. It is therefore natural that the S-matrix o f the Liouville system is trivial. (ii) The energy o f a generally covariant system must be expressed b y a surface term, following the standard method [11 ]. It is . s The action (18) has previously appeared in the literature [ 10] where, however, it was only used as a formal device for deriving the improved energy momentum tensor.
22 November 1984
easily seen that the only possible (t'mite) value o f the energy o f our system is then zero, at least classically. A closely related (but not identical) phenomenon will be discussed, in the context o f a different model, in another paper [12]. (iii) In terms o f ~ and ~ = rr + P , the constraints in the classical theory take the form ~ ± = 2 - 1 ~ 2 - ~ P + 2 ~ - l ' x " - 2 - 1 " X '2 - ~0"X' + A exp(~/~0) and ~gl = X r + ~0, r - ztJ- 7r . From these expressions, we observe that in a recent work [13] b y Banks and Susskind only the special case where P = ~ ' = 0 was treated. F r o m our point o f view, this is not, however, a gauge choice. This explains why these authors had only a collapsed solution for A > 0 (negative scalar curvature). It should also be remarked that with the above restriction we have now [ ~g±(x), ~g±(v)] = 0, which may be interpreted as an indication that the s p a c e - t i m e s swept out b y the deformation generators with this restriction are in fact something different [14] from the ordinary (pseudo-)riemannian space-times. I would like to thank A. Neveu for his comments on the manuscript. [1] A. Polyakov, Phys. Lett. 103B (1981) 207. [2] T. Curtright and C. Thorn, Phys. Rev. Lett. 48 (1982) 1309; E. Braaten, T. Curtright and C. Thorn, Phys. Lett. l18B (1982) 115. [3] J.L. Gervais and A. Neveu, Nucl. Phys. B209 (1982) 125; B224 (1983) 329; CERN preprints TH 3731 andTH 3732. [4] E. D'Hoker and R. Jackiw, Phys. Rev. D26 (1982) 3517. [5] E. D'Hoker, D. Freedman and R. Jackiw, Phys. Rev. D28 (1984) 2583; E. D'Hoker and R. Jackiw, Phys. Rev. Lett. 50 (1983) 1719. [6] S. Fubini, Nuovo Cimento 34A (1976) 521. [7] C. Teitelboim, Phys. Lett. 126B (1983) 41. [8] R. Jackiw, in: Ahead of his time;Bryce S. DeWitt - a collection of essays in honor of his 60th birthday, ed. S. Christensen (Adam Hilger, Bristol, 1983). [9] See, for example, B.S. DeWitt, Phys. Rev. 160 (1967) 1113. [10] R. Marnelius, Nucl. Phys. B211 (1983) 14. [11] T. Regge and C. Teitelboim, Ann. Phys. (NY) 88 (1974) 286. [ 12] T. Yoneya, Higher derivative quantum gravity in two dimensions, CERN preprint TH 3867, to be published in Phys. Lett. B. [13] T. Banks and L. Susskind, Stanford Univ. preprint ITP743 (1983). [ 14] C. Teitelboim, in: General relativity and gravitation: one hundred years after the birth of Albert Einstein, ed. A. Held (Plenum, New York, 1980). 115