#### Charged structure constants from modularity

HJE
Charged structure constants from modularity
Diptarka Das 1
Shouvik Datta 0
Sridip Pal 1
0 Institut fur Theoretische Physik, Eidgenossische Technische Hochschule Zurich
1 Jolla , CA 92093 , U.S.A
We derive a universal formula for the average heavy-heavy-light structure constants for 2d CFTs with non-vanishing u(1) charge. The derivation utilizes the modular properties of one-point functions on the torus. Re nements in N = 2 SCFTs, show that the resulting Cardy-like formula for the structure constants has precisely the same shifts in the central charge as that of the thermodynamic entropy found earlier. This analysis generalizes the recent results by Kraus and Maloney for CFTs with an additional global u(1) symmetry [1]. Our results at large central charge are also shown to match with computations from the holographic dual, which suggest that the averaged CFT three-point coe cient also serves as a useful probe of detecting black hole hair.
Conformal Field Theory; AdS-CFT Correspondence; Black Holes
1 Introduction
2
3
4
5
6
Modular properties of charged CFTs
Derivation of the asymptotic formula
Asymptotics in N = 2 SCFTs
Three-point coe cients from holography
Conclusions
A Properties of Virasoro u(1) algebras B Chern-Simons bulk u(1) gauge eld
in this list include: entanglement entropy (related to minimal surfaces in the bulk) [14]
and conformal blocks (related to geodesics/Witten diagrams) [15{18], to name a few.
A very recent development along these lines, has been to derive the averaged three
point coe cient, i.e. of two heavy and one light operator, using modular properties of
{ 1 {
one-point functions on the torus [1]. In a sense, this goes beyond the Cardy formula, since
the partition function itself is given by the one-point function of the identity operator. In
this paper, we generalize the analysis of [1] to the case when the CFT has a global u(1)
symmetry and a non-zero chemical potential for the same. That is, we are interested in
theories whose partition function has a u(1) grading. In other words, this is the grand
canonical partition function, given by
Z( ; ) = Tr hq L0 c=24yJ0 q L0 c=24yJ0 i :
Owing to the presence of chemical potentials, such partition functions transform as weak
Jacobi forms. We shall make use of these modular transformation properties to attain an
HJEP1(207)83
analogous formula for the three-point coe cients.
Most of our results cover the large class of CFTs which contain Virasoro
u(1)
KacMoody as their chiral algebra. These include, most importantly N = 2 SCFTs and also
theories with W1+1 symmetry. We shall make very few assumputions on the spectrum;
however, our results are valid for any value of the central charge. We also provide
further re nements of the formula for three-point coe cients of three primaries in N = 2
SCFTs. The resulting expression displays precise shifts in the central charge which has
been observed earlier in the exact formula for thermal entropy [19].
The three-point coe cients also admit a description in terms of geodesics in the AdS
dual. The dual background for our CFT setup is that of the BTZ black hole charged
with u(1) hair. This is a solution of the Einstein-Chern-Simons theory. We shall be able
to reproduce the structure constants by evaluating the length of an appropriate geodesic
network. Furthermore, since the black hole is charged, we shall see that bulk Wilson loops
are also necessary in order to reproduce CFT result in its entirety. It has been discussed
in [1] that these CFT results combined with the matching from holographic calculations in
the black hole geometry hints at the notion of the black hole geometry being an averaged
version of heavy microstates. The result of the present work therefore extends this picture
to case when the black hole has additional charged hairs.
Summary of main results. For the convenience of the readers, we provide the main
ndings of our analysis here. We are interested in the three-point coe cient of an uncharged
light operator O with two other heavy operators (denoted by
) with conformal dimension
and charge (Q; Q). The mean value of this quantity shall be denoted by C yO
and is
de ned by the ratio of the spectral density weighted with the 3-point coe cient and the
spectral density itself. In the limit
relates the level to the central charge by k = c=3. It is useful to contrast C yO
with the
uncharged result of [1]
charged primaries, the charged structure constants also possess additional phase which
is of the form e 2 i(q Q q Q)=k. The `holographic' large c limit (c
this formula shall be reproduced from AdS3 gravity; the dual background being that of a
charged BTZ black hole.1
It is also possible to make further re nements to the above formula, for the case of
three-point coe cient of three primaries in N = 2 SCFTs. The derivation uses asymptotic
properties of N = 2 torus blocks. The resultant expression for the average three point
function has the central charge shifted to c 7! c
3. This precise shift has been observed
earlier in [19] for the Cardy formula for the entropy in N = 2 theories and can be
holographically understood as one-loop renormalization of the e ective central charge due to
dressing caused by gravitons, gravitini and the spin-1 gauge eld. This is analogous to the
shift to the shift c 7! c
1, for the non-supersymmetric case, derived in [1].
Outline.
This paper is organised as follows. In section 2 we discuss the relevant modular
transformation properties of one-point functions in presence of u(1) charge. Section 3
contains the derivation of the asymptotics of the charged structure constants. We specialize
to N
= 2 SCFTs in section 4. The large central charge limit of the CFT results are
reproduced from gravity in section 5. We conclude in section 6. Appendices A and B
provide some of the technical details of our analysis. We discuss the relevance of Tauberian
theorems which is relevant to our present context in appendix C.
12
c
; c
Modular properties of charged CFTs
For a 2d CFT with a global abelian symmetry, the torus partition function in the grand
canonical ensemble (i.e. for the non-vanishing chemical potential, , for the u(1) current)
can be written as,
Z( ; ) = Tr hq L0 2c4 q L0 2c4 yJ0 yJ0 i :
(2.1)
Here, the nome and the fugacity are given by q = e2 i and y = e2 i respectively.2 J0 is
the zero-mode of the u(1) current. The modular parameter of the torus given in terms of
the inverse temperature
and the circumference of the spatial circle, L
The modular transformation on
takes real
to imaginary
and vice versa. Therefore,
is chosen to be complex
R + i I to keep things general. This means that the chemical
1In a related context, heavy light conformal blocks in presence of u(1) charge has been studied in [20].
2It is rather unfortunate that the usual q will be used to denote the u(1) charge in due course.
= i =L:
{ 3 {
potentials for both the u(1) charge (J0 + J0) and the fermion number (J0
J0) are turned
on. It is known that under a modular transformation,
transforms as [21, 22]3
= ac ++db , the partition function
Z( ; ) 7!
Z
a + b
;
c + d c + d
= exp
ic k 2
c + d
ic k 2
c + d
Z( ; ):
(2.2)
Here, k is the level of the u(1) Kac-Moody algebra. We also need to consider how a primary
operator, O, transforms under modular transformations. The operator O is located at the
complex elliptic coordinates of the torus, w and w. Under the modular group PSL(2; Z),
HJEP1(207)83
the elliptic coordinate transforms as [24]
w =
w
c + d
The one-point function of the operator O on the torus is given by
O(w; w)j =
w) h
Combining (2.2) and (2.4), we are led to the following transformation property for one-point
function of primaries on the torus (see also [25])
hO(
w;
w)i ; = (c + d)h(c + d)h exp
ic k 2
c + d
ic k 2
c + d
hO(w; w)i ac ++db ; c +d
: (2.6)
This property will play a key role in determining the asymptotics of the charged structure
constants. Note that, the one-point function does not have position dependence, owing
to translation symmetry on the torus.
Moreover, the chiral half of the transformation formula (2.6) is that of a weakly holomorphic Jacobi form of weight h and index k [26].
3
Derivation of the asymptotic formula
We start with the one-point function of a neutral primary on the torus with modular
parameter,
= i =L and chemical potential
=
R + i I . The trace in equation (2.5) can
be rewritten as sum of states as
hOi
=
X
h
jOj
i q
c
24 q
2c4 yQ y
Q :
(3.1)
3See also appendix A of [23] for an example.
{ 4 {
where, the state j
i has the following eigenvalues
(L0 + L0)j
i =
j
i
;
has its advantages; namely, the one-point function of the identity operator is then the
partition function itself. In the low temperature regime, =L ! 1, using the q -expansion
of the sum above, we can write
hOilow-temp ' h jOj ie 2 =L(
1c2 )e2 i Rq e 2 I q+ +
:
= 04 and we have de ned q
= q
is the lightest primary with c yO 6
terms of the charges of . The following analysis is also valid even when
is neutral,
i.e. q
= q
= 0. However, more interestingly one can indeed
nd theories in which the
lightest state is charged. For example in N = 2 super-Virasoro minimal models, the lightest
state, , has the quantum numbers
h =
1
2(k + 2)
;
q =
1
k + 2
:
In this speci c case, is additionally BPS, satisfying h = q=2.
The light operators and y carry opposite u(1) charges. The presence of the terms in the OPE
y
I + O +
are consistent with charge conservation, i.e. the operators appearing on the right hand side
are neutral under the u(1). We also require that
falls within the bound of [21].
From the low temperature expansion (3.3), we can get the expansion in the high
temperature regime by performing the S-modular transformation. This is tantamount to
choosing a = d = 0; c =
equations (2.6) and (3.3) we have,
b = 1,5 which takes
1= and
!
= . Hence, using
hOihigh-temp ' i SO
h jOj ie
2 Lk ( 2R I
2) L
2 L (
1c2 )e 2 Lq+ R e 2 iLq
I
: (3.6)
Let us now rewrite the summation over states in the grand canonical expectation value (3.1)
!
O
e
(3.3)
q in
(3.4)
(3.5)
as an integral
hOi =
Z
d dQ+dQ
three-point coe cient | and is de ned as
5This can also be achieved via b =
is non-degenerate and c yO
is not exponential in the conformal weight.
c = 1. However, this results in a di erent representation of O under
the modular group, PSL(2; Z).
{ 5 {
where, ( ; Q ) is the density of states at conformal dimension
and charges Q
the `unweighted spectral density')
In the above equations and the ones to follow, (Q
Q ; ) refers to the product of the
-functions (Q+
Q ;+) (Q
Q ; ). The notion of the weighted spectral density also
appears in the discussion of OPE convergence in [27]. However, the discussion there deals
with the spectral density weighted by the square of the OPE coe cients.
The weighted spectral density (3.8) also de nes the average three point function,
C yO , the key object we are interested in this paper.
HJEP1(207)83
C yO
followed in [1]. However, now the inversions involve one inverse Fourier transform and two
inverse Laplace transforms, (due to the additional chemical potentials present)
We also note that it is clear that in =L ! 0 regime, the integral (3.7) will be dominated
by large
. Since we are interested in the asymptotic behavior of three point coe cients of
the type heavy-light-heavy, we should then use the high temperature expansion of the torus
grand canonical one point function. This is given by equation (3.6) which we substitute
in (3.11). It turns out that the R and I integrals can be done explicitly and the results
are as follows:
Z 1
d R e
1
from equations (3.12) and (3.11), which leave us with the following integral.
integral can hence be evaluated using the saddle point approximation. The saddle point is
{ 6 {
(3.9)
(3.10)
(3.11)
(3.12)
located at,
s = L
s c
12
O shifts the saddle in the sub-leading order in the 1=
expansion. This
saddle needs to be real and consistent with the high temperature expansion. The relevant
condition (to investigate the large
asymptotics) for the reality of the saddle is
Q2+ + Q2
4k
>
c
12
;
q+2 + q2
4k
<
c
12
The above inequalities have a satisfying interpretation in the bulk dual which we shall
consider in section 5. The zero-mass BTZ black hole is given by c=12. The rst inequality
implies that the asymptotic energy regime is above the charged BTZ black hole threshold,
whilst the scalar corresponding to the light operator
is below the same.6
can therefore
be considered as a perturbative bulk scalar or a massive point particle.
We shall now adopt the following de nitions for clarity of our expressions
These quantities are none other than the spectral ow invariants in the charged CFT. See
appendix A for a short review of the relevant details. Including quadratic
uctuations
around the saddle point, we have,
The nal expression for TO( ; Q ) then reads
One can also carry out the derivation for the density of states for the charged case retaining
the quadratic saddle
uctuations. This is the same calculation as the above with the
operator O being the vacuum/identity. This boils down to setting
;q = 0 =
I =
Q,
Q = 0 = q and using the normalization convention cIII = 1. The density of states then reads
6This suggests the existence of light charged particles in AdS3 in the spirit of the Weak Gravity
Conjecture [28, 29].
c
12
5=4
{ 7 {
c
12
1
; c
1
We can get the asymptotics of the mean charged three point coe cient from (3.8) in the
limit
! 1 and respecting inequalities (3.15). From the de nition of C yO
in terms of
the ratio given in (3.10), this is
C yO
i SO
exp
c
12
This is the central result of our work. The quadratic shifts by the u(1) charges are
consistent with expectations from spectral ow. It is also worthwhile to note that the expression
above exhibits a phase, e 2ki (q Q q Q), due to the presence of the u(1) charge.
In the holographic (large c) limit, i.e, c
O the above simpli es to,
C yO
NO
12
c
c yO e 2 (
Here, NO absorbs the piece independent of ; Q; Q. In section 5, this expression will be
reproduced from the AdS3 dual.
It deserves mention that one can alternatively state and derive the Cardy formula [12]
for density of states, the formula for average heavy-heavy-light coe cient [1] and the
formula (3.19) presented in this work, using the mathematical machinery of Tauberian
theory [30]. This line of approach has also been utilized in [27] in the context of OPE
convergence. We refer the reader to appendix C for a brief review of Tauberian theory, followed
by a discussion of its relevance and usefulness in the present context.
4
Asymptotics in N
= 2 SCFTs
The u(1) current can be naturally embedded as the R-symmetry of the N = 2
superconformal algbera (see appendix A). We shall focus on the three point coe cient of three
primary operators in N = 2 SCFTs. This will enable us to provide further re nements
to the asymptotic formula (3.19). The notion of `heavy' operators is not clear for CFTs
with small values of central charge and the analysis of this section does not include the
N = 2 super-Virasoro minimal models (with central charge c = k3+k2 ). Since, these are
the only possible N = 2 SCFTs with c
complementary to this range.
3 [31], we shall be concerned only with SCFTs { 8 {
The one-point function on the torus (of an operator O with weights (H; H)) can be
expanded in terms of torus conformal blocks (F H (q ; y); F H (q ; y)) as follows.7
hOi =
In the second equality the sum is over all primaries, labelled by . The above equation
generalizes the notion of the partition function (= one-point function of the identity) which
HJEP1(207)83
can be written as a sum of characters. For future convenience, we also rewrite the (4.1) as
an integral (similar to the previous section)
p
dQ+ dQ TO( ; Q ) e
(
1c2 )e2 i RQ e 2 I Q+ F H (q ; y)F H (q ; y); (4.2)
hOi =
Z
d
p
where, the weighted spectral density for primaries, TOp( ), is given by
X
i
TO( ; Q+; Q ) =
h ijOj ii (
i) (Q+
Qi;+) (Q
Qi; ):
(4.3)
Here, we have resorted to the same de nition of the states
i as in (3.2). Since they are
primaries in this context, they additionally obey Lnj ii = 0 for n > 0.
For the one-point function of the identity operator (H = 0), the torus blocks are simply
given by the non-degenerate characters [32]
F
0 (q ; y) = q
(q ; y) = q
+ 2c4 y q
2c4 yq
n=1
(q ; y):
Y1 (1 + yq n 1=2)(1 + y 1q n 1=2)
(1
q n)2
= q
c243 yq #3(( )j3 ) :
(4.4)
The coe cients of the q ; y-series count the number of descendant states and their charge
at each level.8 Note that the overall factor of q
has been accounted for in (4.1).
c=24yq , which cancels out in F
0 (q ; y),
It has been noted in [1] that in the high-energy regime
! 1 and
j log qj2
1,
the torus block for general H is dominated by the character itself. This has been explicitly
shown to be true for Virasoro blocks in [1]; however, it is expected to be true rather
generically. That is, when the intermediate state j i is heavy, the insertion of the light
operator O is a small perturbation to the degenerate character.
F
H (q ; y) = F
0 (q ; y) 1 + O(
1) :
(4.5)
7Recall that the u(1) charge of O is 0. Hence, it su ces to label the torus blocks by (H; H) which are
the only non-vanishing quantum numbers.
8Note that (4.4) is the partition function of a theory of a complex boson and a complex fermion with
central charge equaling 3, barring overall factors of q#. This is the contribution from the descendants of
the N = 2 superconformal algebra.
F
H (q ; y)F H (q ; y)
p
In the nal step we have retained the leading high temperature behaviour. In terms of ,
the product of the holomorphic and anti-holomorphic torus blocks is9
Substituting this in the one-point function (4.2), we have the following expression for the
weighted spectral density in terms of the integral transforms (in the limit
! 1)
2 L
( 2 + 2)L
Here hOi is given, as before, in terms of the dominant contribution from the lightest charged
primary (and a modular transformation thereof)
hOi ;
h jOj ij j
2hO exp
This leads to the same saddle equation as that of (3.13). However, the central charge shifts
consistently all throughout the expression as
c 7! c
3
k 7! k
1:
With the shifts (4.10), the calculation of the mean structure constant proceeds exactly in
the same manner as that of the previous section. The density of states (3.18) also changes
with the same shifts in the central charge. The re ned expression for the average value of
the three-point coe cient of heavy-heavy-light primaries is then given by
This precise shift of the central charge has also been observed in the entropy for N = 2
theories [19]. The non-supersymmetric theories on the other hand show a shift of c to
c
1 [1, 33]. Although we shall not attempt to do here, it would be interesting to recover
this shift from holography as well. It is tempting to speculate such a shift would be caused
by the dressing of the scalars by the gravitons, gravitini and Chern-Simons u(1) gauge
elds, thereby leading to a renormalization of the central charge.
5
Three-point coe cients from holography
In this section we shall reproduce the 3-point coe cient, in the limit of large central
charge (3.20) from holography. The holographic dual to the thermal state of the CFT at
a non-zero chemical potential is that of a BTZ black hole, with additional u(1) hair (since
we are interested in the energy regime above the charged BTZ threshold (3.15)). A u(1)
Chern-Simons gauge eld is also present in the bulk which is dual to the spin-1 conserved
current of the CFT.
The charged BTZ black hole metric is exactly the same as that of the uncharged one;
however, the relation between the mass of black hole and the energy of the CFT gets
modi ed by the non-vanishing charge of the heavy state. For the non-rotating black hole
the metric is given by
mO
m
O
;q =
2k
:
ds2 =
(r2
r+2)dt2 +
r2
dr2
r
+
2 + r2d 2:
In units of `AdS = 1, the radius of the horizon, r+ is given by,
r+ =
where L and L are the shifted zero modes of the holographic stress tensor | see
equation (B.8) of appendix B. Note that r+ is real when the condition (3.15) is satis ed. The
at connections of the u(1) Chern-Simons gauge eld have the following non-vanishing
components
Aw =
Q
k
;
Aw =
Q
k
:
Details on the derivation of the above expressions using the standard GKPW prescription
are provided in appendix B.
We shall reproduce the large c limit of the average three-point coe cient, C yO , given
by equation (3.20). We restrict ourselves to the regime in which the primaries O and
have
by
dual to
1c2 . We shall denote the bulk neutral scalar dual to the CFT probe O
O. It has a mass mO which we take to be large. Similarly
is the charged scalar
of mass m , which is also large. In the large mass (bulk) limit the AdS/CFT
prescription gives us,
(5.1)
(5.2)
(5.3)
(5.4)
coe cient in the constant time slice of the charged BTZ background.
We can use the geodesic approximation in the limit 1
O
;
;q
since we are interested in
Q !
1 limit, by (5.2) r+ ! 1 as well. For
1c2 . Also note that
O, which is the
bulk-to-boundary scalar, the leading amplitude is simply given in terms of its regularized
length (L
log( =r+))
emOL
' const:
rmO
+
This neutral scalar O can be thought of as arising from the fusion of two oppositely charged
scalars,
and y , that wrap the horizon. This vertex is proportional to h jOj i. The
amplitude of the charged scalar
wrapping the horizon in presence of the background
CS eld has two parts, as evident from the relevant classical Euclidean action of a charged
particle of mass m
and charge (q; q) in the curved background:
Sg(eo)desic = Sg(ra)v + SC(S)
Sg(ra)v = m
Z
d
r
g
dx dx
d
d
SC(S) = iq
I
A dx
iq
A dx :
I
(5.5)
(5.6)
(5.7)
Here A ; A
are the dual bulk Chern-Simons elds. The above integrals are along the
closed loop around the horizon as shown in
gure 1. The gravitational on-shell action
simply picks up the horizon area.
Sg(ra)vjon-shell = 2 m r+:
Using the explicit solutions (B.6) for the gauge elds we can next evaluate the on-shell
Wilson loop terms in SC(S),
SC(S)jon-shell = iq
Z 2
0
Putting things together, the amplitude for the geodesic con guration is
h jOj i r+mO e 2 r+m e 2ki (q Q q Q):
If we use the identi cations (5.2) and (5.4) we recover the large c CFT result (3.20) for
C yO , upto the overall normalization.
Z 2
0
2 i
k
A d
iq
A d
=
q Q
q Q :
(5.8)
;q is of the order of c, we can proceed following the arguments as presented in [1].
The massive point particle
backreacts to give rise to a geometry with conical defect.
is the de cit angle and related to mass m . Now, in charged case,
;q (rather
than E , as is done for the uncharged case in [1]) has to be identi ed with the ADM mass,
measured at in nity. Thus we obtain
m
=
1
c
6
r
1
12
;q
c
(5.10)
Plugging in the value of modi ed m leads to the more general result we obtained for CFT (3.19). The CFT result can also be reproduced using Witten diagrams, as has been done in [1].
One needs to make appropriate replacements of the conformal dimensions by their
spectral ow invariant analogues. The additional phase naturally appears in the bulk-to-bulk
propagator of the charged scalar eld
due to the presence of the gauge eld in the bulk.
We remark that the bulk computation can also be reformulated in the Chern-Simons
description of 3d gravity based two copies of the gauge group sl(2; R)
u(1). The
appropriate gauge connections for the charged BTZ black hole can be written down. Such
black holes have already been constructed for the higher-spin charges (in addition to u(1)
charge) in [34]. The relevant background for the present situation can be recovered by
setting these higher-spin charges to zero. The 3-point function can be estimated by using a
network of Wilson lines shown in
gure 1 [35{39]. The Wilson loop (in the representation
R) corresponding to the charged scalar,
, is
log WR(C) = TrR(2 ( x
x)P0):
Since the scalar has mass and charge, we choose P0
L0 + J0. Here, x and x are the
eigenvalues of the ax and ax. (The gauge connections are of the form A = b 1ab + b 1db.)
Evaluating the above Wilson line reproduces (5.8). The radial Wilson line attached to the
boundary (of the uncharged scalar
O) can be incorporated using techniques in [40]. This
requires the Wilson line to be contracted with a chosen state which mimics the insertion of
a primary operator in the CFT. At the trivalent vertex, the Wilson lines are joined using
a suitable intertwiner, which is consistent with the CFT fusion rule (3.5).
6
Conclusions
In this paper, we have been able to extract the mean of the heavy-heavy-light coe cient
for CFTs carrying an u(1) chemical potential conjugate to the conserved current. Our
derivation has relied on modular properties of torus one-point functions in presence of u(1)
chemical potentials. This analysis, therefore, generalizes the one in [1] if chemical
potentials corresponding to additional conserved charges are turned on. The main result (3.19)
contains the same structure as that of one of the uncharged result of [1], with conformal
dimensions replaced by their spectral ow invariants. In addition, there is an extra phase
present owing to the u(1) charge. These modi cations of the uncharged result are clearly
which one would have expected and it is reassuring to see these explicitly. Furthermore, the
mean 3-point coe cient can also be recovered from the bulk dual. This shows that black
holes with additional hairs can be suitably described as an averaged rendition of heavy
microstates, which carry the additional quantum numbers corresponding to the charge.
It would be tantalizing to understand the applicability the above results and its further
generalizations. For instance, the results here can be straightforwardly generalized to
the case when chemical potentials auxiliary to more spin-1 currents from a non-Abelian
Kac-Moody algebra are turned on. Another natural direction would be investigate the
incarnation of the story here, if higher-spin chemical potentials are turned on. Indeed,
the Cardy formula is known perturbatively in powers of the spin-3 chemical potential [22].
One would like to see what happens for the torus one-point function in such a case. This
would also require the knowledge of the one-point functions under modular transformations.
Moreover, one may hope that the average 3-point coe cient can be possibly reproduced
by a suitable network of Wilson lines in the bulk [39].
The formula (3.19) is valid for all ranges of central charge. It will hold true for all
CFTs as long as the required assumptions about the spectrum are made. These results
may have a broader applicability in systems where additional u(1) currents are present
e.g. Luttinger liquids.
A major motivation behind this work arose from studying the Eigenstate
Thermalization Hypothesis in CFTs. It was found in [41] that the heavy-heavy-light coe cient plays a
role in non-universal deviations of the thermal reduced density matrix and its hypothesized
approximation in terms of a single heavy eigenstate. Clearly, a better understanding of ETH
would require an investigation in terms of the Generalized Gibbs Ensemble (GGE), which
has chemical potentials for all conserved quantities turned on. It is therefore necessary to
know the heavy-heavy-light coe cients in presence of additional chemical potentials. The
result in this paper treats the simple possible case and thereby provides a small step in
that direction.
In the context of holography, the three point function contains information of bulk
interaction of scalar elds. This requires bulk probes like the one considered here which
are sensitive to the entire spacetime geometry. In this line of thinking, entanglement has
been used as a tool to reconstruct the bulk geometry, by utilizing its formulation in terms of
the Ryu-Takayanagi surface. However entanglement entropy is a highly dynamic quantity
and is susceptible to UV divergences. In contrast, the mean structure constant is not UV
divergent. The in nities appearing in the lengths of the bulk geodesics which determine
the three point function can thus be unambiguously removed. It would be interesting to
explore this holographic relationship further and to see to what extent bulk geometry can
be constrained using the average OPE coe cient.
Finally, the growth of spectral density has played a crucial role in determining whether
CFTs fall within the `universality class' of being holographic. In order to admit a stringy
dual, CFTs should have a sparse light spectrum and a Hagedorn growth in the density
of primaries at high energies [42{44]. For N = 2 SCFTs, the properties of the elliptic
genus also serve as a diagnostic for determining whether they can admit a putative gravity
dual [45, 46]. It is an exciting avenue to explore the mean values of the OPE coe cient
further and turn their behaviours both at the heavy and light regimes into constraints for
holographic CFTs. We hope that recently developed techniques from higher genus modular
bootstrap will prove to useful in this context [10, 33, 47].
Acknowledgments
It is a pleasure to thank Alejandra Castro, Matthias Gaberdiel, Suresh Govindarajan,
Shamit Kachru, Per Kraus and Alex Maloney for discussions and comments. SD thanks
the organizers and participants of the Summer Workshop at the Aspen Center for Physics,
which is supported National Science Foundation grant PHY-160761. SD thanks the
Department of Physics of UCSD for hospitality and an opportunity to present this work.
The research of SD is supported by the NCCR SwissMap, funded by the Swiss National
Science Foundation. SP thanks the organizers, participants and lecturers of TASI 2017.
DD and SP acknowledge the support provided by the US Department of Energy (DOE)
under cooperative research agreement DE-SC0009919.
A
Properties of Virasoro u(1) algebras
The commutation relations of the Virasoro u(1) generators are given by
[Lm; Ln] = (m
n)Lm+n +
m(m2
1) m+n;0
[Jm; Jn] =
m m+n;0;
[Lm; Jn] =
nJm+n:
c
12
c
3
T (z)T (w)
J (z)J (w)
T (z)J (w)
(z
(z
(z
c=2
c=3
J (w)
w)4 +
(z
w)2
;
w)2 +
(z
w)
:
2T (w)
w)2 +
(z
w)
;
The OPEs involving T and J are (A.1) (A.2) (A.3)
(A.4)
It is well known that this algebra enjoys a spectral ow automorphism | under this
the J0 and L0 modes speci cally transform as (k = c=3)
L00 = L0 + L0 = L0 + J0 +
J00 = J0 +
;
c 2
6
c
3
(see (B.21-23) of [48]). We can now tune to kill the eigenvalue of the J0 charge, J . This
happens at
Consequently, the J0 and L0 eigenvalues become
h0 = h
J 0 = 0:
=
3
c J :
3
2
2c J ;
One can then work with these new charges for which the partition function is
Hence, the spectrum gets shifted by the square of the charge on both on the holomorphic
and anti-holomorphic sides.
relations, given by
We remark that the N = 2 superconformal algebra (A.1) has additional commutation
m
2
We shall restrict our attention to CFTs having just Virasoro
u(1)Kac-Moody as their chiral algebra. In order to obtain the unitarity constraint, we consider
the Kac determinant at level 1. For a given highest weight state jh; qi, characterized by
the conformal dimension h and charge q, the descendant states at level one are L 1jh; qi
and J 1jh; qi. The Gram matrix of these states is
M
(1) =
hh; qjJ1 !
hh; qjL1
J 1jh; qi L 1jh; qi
=
k
q 2h
q
!
:
In a unitary CFT, the Kac determinant should be non-negative. This gives the constraint
det[M
(1)] = 2kh
q
2
0
=)
h
Chern-Simons bulk u(1) gauge eld
The Chern-Simons action with boundary term is given by
SCS =
ik Z
4
M
A ^ dA
A ^ dA
k Z
8
A A + h
A A
:
(B.1)
The normalisation is chosen so that k appears as a level in the current algebra i.e [Jm; Jn] =
km m+n;0. Using the AdS/CFT prescription [49, 50], we obtain the CFT current from the
on-shell boundary variation of the bulk action [51]
SCS =
i Z
2
J w
Aw
where we use complex Euclidean coordinates w =
+ it and w =
it. Varying (B.1)
and comparing with the above equation, we arrive at
Jw =
2
1 J w = {kAw
Jw =
2
1 J w =
{kAw:
It is important to realize that one can either vary Aw or Aw, but not both. In our case,
we are following the convention of [51] and varying Aw, this leads to Jw = 0. Similarly,
(A.7)
(A.8)
(B.2)
(B.3)
J0 =
J0 =
I
dw
I
1
2 { Jw =
1
dw
2 { Jw =
I
dw
I
k
2
dw
k
2
Aw = kAw;
Aw = kAw:
Aw =
Q
k
;
Aw =
Q
k
:
Twgawuge =
A2w =
Q2
4 k Twgawuge =
k
4
A2w =
k
4
I
I
L0 =
L0 =
dwTww =
dwTww =
I
I
dw
dw
Q2
4 k
Q2
4 k
=
=
This implies that the components of the u(1) u(1) gauge eld in the bulk are
The components of the stress-tensor can be obtained by variation of h of (B.1). We have
Now, this contributes to shift in L0 and L0 in following manner, leading to spectral ow
invariant combinations, equation (A.3).
for the anti-holomorphic cases, we are varying Aw whilst keeping Aw
from eq. (B.3), we have
Q2
4 k
Q2
2k
Q2
2k
(B.4)
(B.5)
(B.6)
The relation between asymptotic form of a function and its Laplace transform falls naturally
under the umbrella of a broad class of theorems known as Tauberian Theorems. This branch
of mathematics deals with de ning in nite sums, which are otherwise not summable in the
usual sense (i.e. the partial sum upto nth term does not converge as n ! 1). In Tauberian
theory, one forms a hierarchy of the notion of summability. As one goes further up the
hierarchy, one can sum series which are not summable in the lower hierarchy. The following
example (see chapter 1 of [30] for a detailed introduction) elucidates the scenario.
Consider the sum P
k=0( 1)k(k + 1). Evidently, this is not summable in the normal
sense. But one can de ne
f ( )
X( 1)k(k + 1)e n
For
> 0, however, this is summable in normal sense and we nd
k=0
e
2
(e + 1)2
It's easy to see, e2 = e + 1 2 is well de ned at
say f ( ) goes to 1=4 as
goes to 0. Now one de nes,
= 0 and equals to 1=4. Thus, one can
X( 1)k(k + 1)
k=0
=
New no|t{iozn}of sum
1
4
:
This notion of sum is called Abel sum. It's easy to see that if a series is summable in
normal sense, it is summable in Abel sense, but not the other way around. The Tauberian
theorems specify the conditions under which the higher notion of summability (e.g. Abel
summability) implies the lower notion of summability (e.g. normal summability). In its
generalized version, one can deduce the asymptotic behavior of normal sum if one knows
the asymptotic behavior of Abel sum. At this point, it deserves mention that the continuous
version of Abel sum is precisely the Laplace transform.
The notion of Abel sum (or Laplace transform as its continuous avatar) becomes
relevant, in the present context, since we are looking at quantities of the form
HJEP1(207)83
and trying to nd out the behavior of P
k=0 ak asymptotically. To make the analogy more
precise, we note that the partition function of a CFT on a torus is given by,
Z
Z( ) =
d
( ) e (
c=12) ;
( ) =
X
k
(
k);
where
is the inverse temperature and length of one of the cycles of torus. Now, knowing
the form of Z( ) as
goes to 0 enables us to deduce the asymptotic form of Rc=12 d 0 ( 0)
as
goes to 1. The power of Tauberian theory comes from the fact that one does not have
to impose any regularity condition on ( ). It can be seen that deducing the Cardy formula
for density of states of a CFT is equivalent to the following theorem (for more details, see
Theorem 21.1 (chapter 4) and its immediate application in Example 21.2 of [30]).
Theorem 1 Let S(v) be a non decreasing function and S(v) = 0 for v < 0. Let F ( ) be
(C.4)
(C.5)
(C.6)
(C.7)
S(v)
S0(v) =
evh(v)+f(h(v))
h(v)p2 f 00(h(v))
;
as v ! 1
(C.8)
and
Then,
Z 1
0
Z 1
0
F ( ) =
e
v dS(v) =
S(v)e
vdv;
for Re( ) > 0;
e f( )F ( ) ! 1;
as
! 0
uniformly in every angle jarg( )j
0 < 2
conditions as
2 R and
& 0, in which we have 0 < ( )
2
,
. Furthermore, if f ( ) satis es the following
f ( ) is real and positive,
f 0( ) % 1,
jf0( )j
pf00( ) = O
( ) ,
f 00( + z) = O (f 00( )) uniformly for jzj
( ).
To derive the Cardy formula [12], we make following identi cations10
F ( ) = Z( );
and concretely, this yields
Now, one can argue that heavy excited states lie very closely to each other so that
density of states becomes almost continuous, which allows us to take a derivative with
and deduce the Cardy formula [12] for density of states of a CFT
( )
1
p
!1
2 12
c 1=4
c
12
3=4
exp 4
r c
12
c
12
Similarly, one can justify the result of [1] by appealing to same theorem provided the
quantity estimated i.e. Rc=12 d 0 TO( 0) is a non decreasing function of . This requirement
can be however relaxed by observing that if all the c yO
comes with same phase, one can
then de ne a new quantity T~O by absorbing the phase. This ensures that Rc=12 d 0 T~O( 0)
is a non decreasing function of
. The more general scenario without this caveat can be
explored by using the Tauberian theory of more general function S(v). For now, with this
caveat in mind, Rc=12 d 0 TO( 0) and Rc=12 d 0 ( 0) are obtained. Once again, going from
integral to integrand requires taking derivative and is justi ed by appealing to the extra
input that at heavy excited states TO becomes a smooth function of . Hence, this justi es
the de nition
C yO
=
TO( )
( )
:
Cc-yaOvg
=
R
R
c=12 d 0 TO( 0)
c=12 d 0 ( 0)
:
(C.11)
(C.12)
(C.13)
It is worthwhile to note that one can de ne following cumulative average as well
In the large limit, the above quantity has exactly the same exponential behavior as
C yO . It deserves mention that in order to apply the above theorem for the OPE
coefcient, one needs to ensure f ( ) is real and positive in the limit
& 0. Since, f ( ) goes
like
In fact,
such that
O log( ) + 4 2
c
12
, the positivity is guaranteed assuming existence of state
< 1c2 . (Note that, log( ) < 0 for
< 1c2 becomes a necessary condition as
< 1 and
O > 0 by unitarity bound.)
f 0( ) > 0 as
& 0 is required
for the above theorem to hold. Furthermore, one can extend the prescription and use the
Tauberian machinery to the charged case which we have studied here. In the charged case,
where
carries charge (q; q), we have
replaced by
;q =
requirement becomes
;q < 1c2 .
21k q2 + q2 and the
10In terms of , the theorem requires that F ( ) = Rc1=12 d
( )e ( c=12). The partition function,
Z( ) = R 1 d
0
( )e (
c=12)), however di ers from F ( ). Nonetheless, in
! 0 limit, the dominant
contribution to Z( ) arises from large . Hence, the fact Z( ) di ers from F ( ), does not a ect this result.
In the
! 0 regime, we indeed have F ( ) = Z( ).
Open Access.
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[1] P. Kraus and A. Maloney, A Cardy formula for three-point coe cients or how the black hole
got its spots, JHEP 05 (2017) 160 [arXiv:1608.03284] [INSPIRE].
HJEP1(207)83
the 3D Ising model with the conformal bootstrap, Phys. Rev. D 86 (2012) 025022
[4] S. Rychkov and Z.M. Tan, The -expansion from conformal eld theory, J. Phys. A 48
[5] S. Hellerman, A universal inequality for CFT and quantum gravity, JHEP 08 (2011) 130
[arXiv:0902.2790] [INSPIRE].
[arXiv:1307.6562] [INSPIRE].
[6] D. Friedan and C.A. Keller, Constraints on 2d CFT partition functions, JHEP 10 (2013) 180
B 270 (1986) 186 [INSPIRE].
3609 [hep-th/9806026] [INSPIRE].
[7] T. Hartman, C.A. Keller and B. Stoica, Universal spectrum of 2d conformal eld theory in
the large c limit, JHEP 09 (2014) 118 [arXiv:1405.5137] [INSPIRE].
[8] S. Collier, Y.-H. Lin and X. Yin, Modular bootstrap revisited, arXiv:1608.06241 [INSPIRE].
[9] S. Collier, P. Kravchuk, Y.-H. Lin and X. Yin, Bootstrapping the spectral function: on the
uniqueness of Liouville and the universality of BTZ, arXiv:1702.00423 [INSPIRE].
[10] M. Cho, S. Collier and X. Yin, Genus two modular bootstrap, arXiv:1705.05865 [INSPIRE].
[11] S. Ribault, Conformal eld theory on the plane, arXiv:1406.4290 [INSPIRE].
[12] J.L. Cardy, Operator content of two-dimensional conformally invariant theories, Nucl. Phys.
[13] S. Carlip, What we don't know about BTZ black hole entropy, Class. Quant. Grav. 15 (1998)
[14] S. Ryu and T. Takayanagi, Holographic derivation of entanglement entropy from AdS/CFT,
Phys. Rev. Lett. 96 (2006) 181602 [hep-th/0603001] [INSPIRE].
[15] A.L. Fitzpatrick, J. Kaplan and M.T. Walters, Universality of long-distance AdS physics
from the CFT bootstrap, JHEP 08 (2014) 145 [arXiv:1403.6829] [INSPIRE].
[16] E. Hijano, P. Kraus, E. Perlmutter and R. Snively, Semiclassical Virasoro blocks from AdS3
gravity, JHEP 12 (2015) 077 [arXiv:1508.04987] [INSPIRE].
[17] E. Hijano, P. Kraus, E. Perlmutter and R. Snively, Witten diagrams revisited: the AdS
geometry of conformal blocks, JHEP 01 (2016) 146 [arXiv:1508.00501] [INSPIRE].
[18] E. Hijano, P. Kraus and R. Snively, Worldline approach to semi-classical conformal blocks,
JHEP 07 (2015) 131 [arXiv:1501.02260] [INSPIRE].
and near-extremal CFTs, JHEP 08 (2016) 023 [arXiv:1603.08524] [INSPIRE].
from classical background elds, JHEP 11 (2015) 200 [arXiv:1501.05315] [INSPIRE].
2d CFT and 3d gravity, JHEP 08 (2016) 041 [arXiv:1603.09745] [INSPIRE].
(2012) 103 [arXiv:1203.0015] [INSPIRE].
[23] J. de Boer and J.I. Jottar, Boundary conditions and partition functions in higher spin
AdS3/CFT2, JHEP 04 (2016) 107 [arXiv:1407.3844] [INSPIRE].
modular forms, arXiv:1208.4074 [INSPIRE].
[26] A. Dabholkar, S. Murthy and D. Zagier, Quantum black holes, wall crossing and mock
[27] D. Pappadopulo, S. Rychkov, J. Espin and R. Rattazzi, OPE convergence in conformal eld
theory, Phys. Rev. D 86 (2012) 105043 [arXiv:1208.6449] [INSPIRE].
[28] N. Arkani-Hamed, L. Motl, A. Nicolis and C. Vafa, The string landscape, black holes and
gravity as the weakest force, JHEP 06 (2007) 060 [hep-th/0601001] [INSPIRE].
[29] M. Montero, G. Shiu and P. Soler, The weak gravity conjecture in three dimensions, JHEP
10 (2016) 159 [arXiv:1606.08438] [INSPIRE].
[30] J. Korevaar, Tauberian theory: a century of developments, Grund. Math. Wiss. 329,
Springer Science & Business Media, Berlin Heidelberg Germany, (2013).
[31] W. Boucher, D. Friedan and A. Kent, Determinant formulae and unitarity for the N = 2
superconformal algebras in two-dimensions or exact results on string compacti cation, Phys.
Lett. B 172 (1986) 316 [INSPIRE].
[32] C.A. Keller and H. Ooguri, Modular constraints on Calabi-Yau compacti cations, Commun.
Math. Phys. 324 (2013) 107 [arXiv:1209.4649] [INSPIRE].
[33] J. Cardy, A. Maloney and H. Max eld, A new handle on three-point coe cients: OPE
asymptotics from genus two modular invariance, JHEP 10 (2017) 136 [arXiv:1705.05855]
[INSPIRE].
[arXiv:1303.1946] [INSPIRE].
04 (2014) 089 [arXiv:1306.4347] [INSPIRE].
[34] S. Datta and J.R. David, Black holes in higher spin supergravity, JHEP 07 (2013) 110
[35] J. de Boer and J.I. Jottar, Entanglement entropy and higher spin holography in AdS3, JHEP
[36] M. Ammon, A. Castro and N. Iqbal, Wilson lines and entanglement entropy in higher spin
gravity, JHEP 10 (2013) 110 [arXiv:1306.4338] [INSPIRE].
theories, JHEP 03 (2015) 124 [arXiv:1410.2870] [INSPIRE].
[37] A. Castro and E. Llabres, Unravelling holographic entanglement entropy in higher spin
[38] E. Hijano and P. Kraus, A new spin on entanglement entropy, JHEP 12 (2014) 041
[arXiv:1406.1804] [INSPIRE].
[arXiv:1412.2759] [INSPIRE].
gravitational open Wilson networks, JHEP 06 (2016) 119 [arXiv:1602.02962] [INSPIRE].
conformal blocks, JHEP 09 (2017) 149 [arXiv:1706.00047] [INSPIRE].
HJEP1(207)83
and 3d gravity, Annales Henri Poincare 17 (2016) 2623 [arXiv:1503.04800] [INSPIRE].
supersymmetric index, JHEP 05 (2016) 158 [arXiv:1512.00010] [INSPIRE].
arXiv:1705.05862 [INSPIRE].
04 (2016) 077 [arXiv:1512.00073] [INSPIRE].
[hep-th/9802150] [INSPIRE].
(2008) 193 [hep-th/0609074] [INSPIRE].