Bounds on OPE coefficients from interference effects in the conformal collider
JHE
Bounds on OPE coefficients from interference effects
Clay Co´rdova 0 2
Juan Maldacena 0 2
Gustavo J. Turiaci 0 1
Jadwin Hall 0
Washington Road 0
Princeton 0
U.S.A. 0
String Theories
0 1 Einstein Drive , Princeton, NJ , U.S.A
1 Physics Department, Princeton University
2 School of Natural Sciences, Institute for Advanced Study
We apply the average null energy condition to obtain upper bounds on the threepoint function coefficients of stress tensors and a scalar operator, hT T Oi, in general CFTs. We also constrain the gravitational anomaly of U(1) currents in fourdimensional CFTs, which are encoded in threepoint functions of the form hT T J i. In theories with a large N AdS dual we translate these bounds into constraints on the coefficient of a higher derivative bulk term of the form R φ W 2. We speculate that these bounds also apply in deSitter. In this case our results constrain inflationary observables, such as the amplitude for chiral gravity waves that originate from higher derivative terms in the Lagrangian of the form φ W W ∗.
AdSCFT Correspondence; Conformal Field Theory; Anomalies in Field and

HJEP1(207)3
1 Introduction
2 ANEC and the conformal collider
2.1
2.2
The average null energy condition
The conformal collider
2.2.1
External states created by T
3 Bounds on T T O in d ≥ 4
3.1
Analysis of the bound
4 Bounds on T T O in d = 3
4.1
ChernSimons matter theories
4.2 3d Ising model
5 Bounds on T T J in d = 4
5.1 Supersymmetry and the Rcurrent
3.2 Free field theories and destructive interference
6 Bounds on coefficients of the AdS effective action
7 Constraints for deSitter and inflation
7.1
Comments on scalartensortensor threepoint functions
A Absence of positive local operators
B Details of the collider calculation
B.1 Normalized states
B.2 Threepoint functions
B.3 Energy matrix
C Free scalar correlators
D hT T Oi parityodd structures in d = 3
E hT T J i threepoint function
F Computing the bound in the gravity theory
F.1 Fourdimensional case
– 1 –
Introduction
In this paper we investigate some implications of the average null energy condition in
conformal field theories. We consider the conformal collider physics experiment discussed in [1].
In that setup, we produce a localized excitation by acting with a smeared operator near the
origin of spacetime. Then we measure the energy flux at infinity per unit angle. Requiring
that the energy flux is positive imposes constraints on the threepoint function coefficients.
This method was used to constrain threepoint functions of the stress tensor in [
1–3
].
In this paper we use this same method to constrain the threepoint functions of two
stress tensors and another operator hT T Oi. The new idea consists of creating the initial
state by a linear combination of a stress tensor operator and the operator O. The
threepoint function hT T Oi appears as a kind of interference term in the expression for the
energy. Requiring that the total contribution to the energy flux is positive imposes a
nontrivial upper bound on the absolute magnitude of this threepoint correlator. We apply
these ideas to general scalar operators O as well as conserved currents with spin one, J ,
where we use it to put bounds on the gravitational anomaly in d = 4 CFTs. Because the
bound arises from quantum mechanical interference effects, these bounds are stronger than
those obtained in states created by a single primary local operator and its descendants
(though the resulting bounds involve more OPE coefficients).
This energy flux at infinity is given by an integral of the stress tensor. On the boundary
of Minkowski space this integral is simply the average null energy E = R dx−T−−. We
review this in section 2. Physically, we expect that this energy should be positive for
all angles. Recently, the average null energy condition was proven using entanglement
entropy methods [4] as well as reflection positivity euclidean methods [5]. When we create
a localized state using the stress tensor, this energy distribution is completely determined
by the threepoint function of the stress tensor.
Two of the insertions correspond to
the insertions creating the state in the bra and the ket. The third corresponds to the
one measuring the energy flux at infinity. The resulting bounds could also be obtained
by requiring standard reflection positivity of the euclidean theory [6, 7]. However, the
conformal collider calculations provide an efficient way to extract the results.
One of our main results is a sum rule constraining the OPE coefficients of scalar
primary operators O with the energymomentum tensor T . In spacetime dimensions d ≥ 4
there is a single OPE coefficient controlling the hT T Oi threepoint function. We find that
this data is constrained as
CT T Oi 2 f (Δi) ≤ NB ,
some simple physical consequences such as its interpretation in free field theories, large N
holographic systems, and general implications for the asymptotics of OPE coefficients.
In section 4 we consider analogous results in spacetime dimension three. This case
is special because the threepoint functions of interest admit both parity preserving and
parity violating structures. The bounds we find generalize those recently obtained in [8].
We apply our results to large N ChernSimons matter theories, and further use them to
obtain predictions on OPE coefficients CT T O for scalars in the Ising model using the recent
results of the conformal bootstrap [9]. For instance, we find that operator ε has an OPE
coefficient constrained as
where the righthand side is the value in the free scalar theory based on the field φ.
In section 5 we consider bounds in fourdimensional CFTs with a global symmetry
current J . We apply the same techniques to obtain universal constraints on the gravitational
anomaly of the current J.
In section 6 we show that the hT T Oi correlator can be generated from a gravity theory
in AdSd+1 through a higher derivative term, R φW 2, in the bulk effective action. We match
the coefficient of this term to the CT T O coefficient in the boundary theory by performing
the same collider experiment in the bulk, where it involves propagation through a shock
wave. One interesting feature of this presentation is that the resulting bound is independent
of the mass of φ. Thus, the Δ dependence of (1.1) is purely kinematic and results from
translating the boundary threepoint function coefficient to a bulk interaction. We use our
AdS presentation to show that α′ corrections satisfy the bound.
In section 7 we extrapolate the bounds we obtained in AdS to “quasi bounds” on the
coefficients of the effective action in de Sitter space. We call them “quasibounds” because,
unfortunately, for deSitter we do not know how to prove a sharp bound. We can think of
these as a good indication for where the bulk effective theory should break down. We apply
these “quasibounds” to constrain the amplitude of chiral gravity waves, and to constrain
the violations of the inflationary “consistency condition” for the twopoint function. Both
of these effect arise from higher curvature couplings of the form φW 2 or φW W ∗.
In the appendices we include more explicit derivations of the material in the main
CT T ε ≤ 1.751CT T :φ2: ,
(1.3)
HJEP1(207)3
sections.
2
2.1
ANEC and the conformal collider
The average null energy condition
The null energy condition is a central assumption in many classical theorems of general
relativity. These results allow us to exclude unphysical spacetimes where causality violation,
naked singularities, or other physical pathologies occur [10].
If we move beyond classical field theory, these results appear to be in doubt. Quantum
effects lead to fluctuations that prohibit any local operator from having a positive
expectation value in every state [11]. (We review these ideas in appendix A.) In particular the
– 3 –
(a)
(b)
i
+
J
+
i
0
HJEP1(207)3
(blue) is measured far away by a calorimeter (red). (b) For a CFT, this is equivalent to measuring
the energy at null infinity J +.
local energy density and other components of the energymomentum tensor have negative
expectation value in some states.
Deeper investigation reveals a potential resolution. While components of the
energymomentum tensor are pointwise nonpositive, a weaker hypothesis, the socalled average
null energy condition, is often sufficient to enforce causal behavior [12]. This condition
states that the integral along a complete null geodesic of the null energy density is a
positive definite operator
E =
Z ∞
−∞
dx− T−− ≥ 0 .
(2.1)
Recently there has been significant interest in understanding the average null energy
condition (2.1) in the context of local quantum field theories. In [5], an argument was given
establishing (2.1) in conformal field theories by examining the constraints of causality on the
lightcone operator product expansion. In [4], an alternative argument was given linking the
average null energy operator to entanglement entropy, then establishing positivity using
strong subadditivity. These information theoretic methods have also been extended to
obtain new inequalities strengthening (2.1) [13].
Given that the average null energy in quantum field theory is now a theorem, it is
interesting to take it as input and use it to constrain conformal field theory data.
2.2
The conformal collider
An efficient way to extract consequences of the average null energy condition in CFTs is
to use the conformal collider setup of [1]. This technique is closely related to deep inelastic
scattering experiments in conformal field theory [1, 14]. As we review, in the context of
AdS/CFT these bounds arise from demanding causality of the bulk theory in a shockwave
background.
The specific physical problem of interest is to create a disturbance in a conformal field
theory and then to measure the correlation of energy deposited at various angles at future
null infinity (see figure 1).
– 4 –
The states in which we measure the energy are obtained by acting with local
opermomentum q.1 Thus we examine the state
ators O(x) on the Lorentzian vacuum 0i. We further give these states definite timelike
Z
O(q, λ)i = N
ddx e−iqt λ · O(x)0i ,
(2.2)
where λ is a polarization tensor accounting for the possible spin of O, and N is a
normalization factor defined such that (2.2) has unit norm.
We now measure the energy at null infinity in this state. In d dimensions null infinity
is a sphere Sd−2 and we parameterize it by a unit vector n.
hE (n)iλ·O = lim rd−2 Z ∞
r→∞
−∞
dx− hO(q, λ)T−−(x−, rn)O(q, λ)i .
(2.3)
The average null energy condition implies that the resulting function is nonnegative as a
function of the direction n.
Since we are working in a conformal field theory this energy expectation value may be
explicitly evaluated. Indeed the object being integrated in (2.3) is a threepoint function
hOT Oi in Lorentzian signature with a prescribed operator ordering. Thus, the result
of (2.3) is an explicit function of OPE coefficients.
2.2.1
External states created by T
Let us review the essential details of this calculation in the case where the external state
is created by an energy momentum tensor. In general in d ≥ 4 spacetime dimensions, the
threepoint function of energymomentum tensors may be parameterized in terms of three
independent coefficients
hT T T i = NBhT T T iB + NF hT T T iF + NV hT T T iV ,
(2.4)
where the various B, F, V structures are those that arise in a theory of respectively free
bosons, fermions, or (d − 2)/2 forms.2 Our conventions are such that for free fields, NB
counts the number of real scalars, NF the total number of fermionic degrees of freedom
(e.g. it is 2⌊d/2⌋ for a Dirac fermion), and NV counts the number of degrees of freedom in
a (d − 2)/2 form (for a single such field this number is Γ(d − 1)/Γ(d/2)2).
A single linear combination of these coefficients is fixed by the conformal Ward identity,
and related to the twopoint function coefficient CT of energy momentum tensors (see
equation (B.7) for our conventions on the twopoint function)
CT =
1
Ω2d−1
d
d − 1
d
2
d
2
2
NB +
NF +
NV
.
(2.5)
1For technical reasons it is sometimes useful to create a localized wavepacket instead of an exact
momentum eigenstate. This subtlety will not affect our discussion.
2In odd d there is no free field associated to the structure parameterized by NV , but nevertheless there
is still a structure. See [
3, 15
] for details.
– 5 –
where Ωn is the area of a sphere Sn.3 As another point of reference let us briefly specialize
to the case of fourdimensional theories. In that case, the coefficients of the threepoint
function are related to conformal anomalies a, c that parameterize the trace of the
energymomentum tensor in a general metric background
hTµ µ i[g] =
c
T ’s is parity invariant.4 It follows that the most general expression for the null energy is
where the constants have been fixed so that the total energy of the state is q, and t2 and
t4 are computable functions of NB, NF , NV .
A useful way to understand the answer is to view the vector n as fixed and to decompose
the states (parameterized by their polarizations) under the remaining symmetry group
SO(d − 2). For example, the polarization that has spin zero under rotations around the ~n
axis is
0
λij ∝
ninj − (d − 1)
δij
In a similar way we can write polarization tensors that have spin one and spin two under
rotation around the ~n axis. The energy flux in the direction n is the same for every state
in a fixed SO(d − 2) representation, and we denote them by qTi/Ωd−2. Explicitly carrying
out the integrals gives:
(2.6)
(2.7)
T0 =
T1 =
t2
t2
1 − d − 1 − d2 − 1
1 − d − 1 − d2 − 1
2t4
2t4
+
d − 2
d − 1 (t2 + t4) = ρ0(d)
NB
CT
,
t2
T2 = 1 − d − 1 − d2 − 1
2t4
= ρ2(d)
NV
CT
where the index labels the SO(d−2) charge and in the above ρi(d) is a positive function that
depends only on the spacetime dimension (and not the OPE coefficients). Their explicit
form is given in equation (B.33).
43IΩnn−d1==32tπhne/2th/Γre(enp/2o)in.t function has a parity odd piece which we discuss in section 4.
– 6 –
T0 = T1 = T2 = 1.
condition implies the inequalities
Additional symmetries imply constraints on the parameters above. In any
superconformal field theory we have t4 = 0. For holographic CFTs dual to Einstein gravity
the parameters are t2 = t4 = 0, giving angle independent energy onepoint functions
Returning to the general discussion, we can see from (2.10) that the average null energy
NB ≥ 0 ,
NF ≥ 0 ,
NV ≥ 0 .
(2.11)
One significant remark concerning the bounds (2.11) is that they may clearly be
saturated in free field theories. Conversely, it has been argued [16] that any theory that
saturates the conformal collider bounds must be free. The fact that the bounds may be
saturated in actual CFTs illustrates that the conformal collider is an efficient way of
extracting the implications of the average null energy condition. Namely, we could not possibly
get a stronger bound, otherwise we would run into a contradiction with free theories.
3
Bounds on T T O in d ≥ 4
We now turn to our main generalization of the conformal collider bounds reviewed in
section 2.2. We explore the consequences of the average null energy condition in more
general states than those created by a single primary operator. Specifically in this section
we will investigate states which are obtained by a linear combination of primary operators.
We will find that the average null energy condition in such states yields new inequalities
on OPE coefficients.
In this section, the states we consider will be created by a linear combination of the
energymomentum tensor and a general scalar hermitian operator O. We parameterize such
a state in terms of normalized coefficients vi
Ψi = v1T (q, λ)i + v2O(q)i .
The energy onepoint function in the collider experiment is now a matrix
hT (q, λ)E (n)T (q, λ)i hT (q, λ)E (n)O(q)i !
hT (q, λ)E (n)O(q)i∗
stronger condition than requiring that the diagonal entries are positive and will imply new
inequalities on OPE coefficients.
The majority of the entries in this matrix have already been computed. For instance, in
section 2.2.1 we reviewed the portion of the matrix involving the energy expectation value
in states created by the energy momentum tensor. Even simpler is the entry involving the
expectation value in the scalar state which gives rise to a uniform energy distribution
– 7 –
(3.1)
(3.2)
(3.3)
It remains to determine the offdiagonal entries in the matrix. It is again useful to
organize the expected answer using the rotation group on the null sphere. Clearly we have
hT (q, λ)E (n)O(q)i ∼ λij ninj .
Therefore, the only polarization of the energy momentum tensor that participates in the
nontrivial interference terms is the scalar T0 aligned along the axis n (see equation (2.9)).
To extract this matrix element we require the threepoint function hT T Oi. In all d ≥ 4,
the conservation constraints on T imply that this correlator is fixed in terms of a single
OPE coefficient CT T O. We set conventions for our normalization of this OPE coefficient
by examining a simple OPE channel. Specifically we restrict all operators to a twoplane,
spanned by complex coordinates z, z¯. Then the OPE is
Tzz(z)Tz¯z¯(0) ∼ zCT2dT−OΔ O(0) .
If we further assume that O is hermitian then the OPE coefficient CT T O is real.
Additional details of this correlator including the full ddimensional Lorentz covariant OPE and
relation to the spinning correlator formalism of [17] are given in appendix B.
Based on these remarks, we can in general parameterize the energy flux in the direction
n coming from the offdiagonal matrix element (3.4) as
where h(Δ) is some universal function that may be extracted from the conformal collider
calculation, and the factors of CT and C
O arise from normalizing the states. The relevant
portion of the energy matrix (3.2) is twobytwo and takes the form
Positivity of this matrix therefore leads to the constraint
q
Ωd−2
T0
√CCT∗TTCOO h(Δ)
T plus a general linear combination of primary scalar operators. Positivity of the resulting
energy matrix is then equivalent to the following sum rule
In appendix B we explicitly compute the function h(Δ) (see equation (B.48)). By combining
the result with the expression (2.10), we may reexpress the bound as
X
– 8 –
We now turn to an analysis of the consequences of the general bound (3.10). The function
f (Δ) has a number of significant properties.5
• Expanded near the unitarity bound we find a first order pole:
as fast as √x.
Therefore in any family of theories, an operator O which is parametrically becoming
free (i.e. Δ = (d − 2)/2 + x with x tending to zero) must have CT T O vanish at least
• For large Δ we find exponential growth
f
d − 2
2
+ x
1
∼ x
.
f (Δ) ∼
4Δ
Δ 72d +4
.
ρ(Δ) CT T O
2
C
O
≤
Δ 72d +3
4Δ
We may use this growth to approximate the sum in the bound for scalar operators of
large Δ. Indeed, let ρ(Δ) denote the asymptotic density of scalar primary operators.
From convergence of the sum we then deduce that for large Δ the spectral weighted
OPE coefficients must decay exponentially fast
These estimates agree with those implied by convergence of the OPE expansion found
in [19] for scalar operators.
• If Δ is an even integer greater than or equal to 2d we find that f (Δ) vanishes. We can
understand the necessity of this as follows. We can imagine a large N CFT dual to
weakly coupled theory of gravity. In such theories we can consider the sequence of
operators O =: T AB∂2nTAB :. At large N the dimensions of these operators are fixed to
Δ = 2d + 2n. Moreover, for these operators CT2 T O is of order C2 . Thus, compatibility
with the bound (3.10) for large CT , requires that f (Δ) vanishes at these locations.
CO
T
The above argument does not explain why f (Δ) has double zeros. But the double
zeros imply that the bound may be obeyed at subleading order, where we include
the anomalous dimensions of these operators which scale as 1/CT , by truncating the
sum on n.6
• The function f (Δ) is nonzero for Δ = d. Therefore the bound (3.10) may be applied
to marginal operators. In that context, it constrains the change in CT at leading
order in conformal perturbation theory.
5A function with similar properties was obtained in [18] in a different context.
6We thank E. Perlmutter for comments on this point.
– 9 –
Free field theories and destructive interference
Let us investigate the bound further in free field theories. These examples are interesting
because the bound (3.10) is saturated.
Consider first a theory of a free real boson φ in dimension d. There is a Z2 global
symmetry under which φ is odd and the energymomentum tensor T is even. Therefore we
need only consider scalars made from an even number of φ’s. Since the explicit expression
for T is quadratic in the free fields, the only possible scalars that may contribute to the
bound are : φ2 : and : φ4 :.
By a simple inspection of the Wick contractions we deduce that : φ4 : has vanishing
T T O correlation function.7 Meanwhile : φ2 : has
CT T O2 =
C
O
(d − 2)4Γ(d/2 + 1)4
8π2d(d − 1)4
This exactly saturates the bound (3.10).
We can also consider the bound applied to free fields of different spin. In d = 4 the
theory of free fermions or free gauge bosons have vanishing NB. Therefore the bound
implies that for all scalar operators O either CT T O vanishes, or O has dimension 2d + 2n
for nonnegative integer n.
that Fµν+ F −µν
It is straightforward to directly verify this prediction. For instance consider the free
vector. The gauge invariant field strength gives rise to two local operators Fµν+ and Fµν− ,
which are respectively selfdual and antiselfdual twoforms.
Note that this free field
theory enjoys a continuous electromagnetic duality symmetry under which Fµν± rotate with
opposite charge. The energymomentum tensor Tµν is neutral under this transformation,
and hence a scalar operator O with nonvanishing CT T O must also be neutral. If we recall
vanishes identically, then we see that the lowest dimension neutral scalar
operator is (Fµν+ F +µν )(Fα−βF −αβ). Since this has dimension eight, the weight function f (Δ)
vanishes. Moreover all other scalar operators that are neutral have larger even integer
dimension. Thus, the bound is obeyed.
A more physical way to understand why the bound is saturated in the free scalar
theory is to visualize the state created by local operators.
Let us consider the action of an operator with nonzero energy but zero spatial
momentum. If the operator is a bilinear in the fields, such as the stress tensor in a free theory,
then it will create a pair of particles with back to back spatial momenta. Of course, the
operator creates a quantum mechanical superposition of states where these momenta point
in various directions. For a scalar bilinear operator we get an swave superposition. For
the stress tensor we get a superposition determined by the polarization tensor.
As in previous sections, we measure the energy in the angular direction n and hence
can focus on the properties of the wavefunction for the pair of particles in that particular
direction. As in section 2.2.1 it is convenient to decompose the polarization tensors of the
operators according to their angular momentum around the n axis. We can then easily
7The contractions imply that hT T : φ4 :i ∝ hT : φ2 :ihT : φ2 :i, which is zero since twopoint functions
of different operators vanish.
+
T0
+
(d)
O
+
1/2
(b)
T1
1/2
1/2
φ2
+
1
(c)
T2
+
(f)
J0
In (a,b,c) we consider a stress tensor operator. We examine the wavefunction along the direction
specified by the long arrow and we decompose the stress tensor according to the spin around that
axis. (a) The spin zero state is obtained for scalars, spin one for fermions (b) and spin two for
vectors or selfdual forms (c). (d) is the state produced by a scalar operator with can interfere with
(a). (e) is produced by a current with spin one along the observation axis and can interfere with
(b). Finally (f) is a current with spin zero along the observation axis in a theory of scalars. It
produces two different real scalars in the back to back configuration and cannot interfere with (a).
check that a spin zero state T0 can be produced only in a theory of scalars, a spin one state
T1 can be produced only in a theory of fermions, and T2 only in a theory of vectors (or
d/2 − 1 forms), see figures 2(a,b,c). This explains formula 2.10.
A scalar operator of the form O =: φ2 :, where φ is an elementary scalar, can also
produce a back to back combination of scalar particles, see figure 2(d). Along the direction
of observation this combination has the same form as the one produced by T0, in figure 2(a).
It is clear that we can make a quantum mechanical superposition so that the wavefunction
for the pair vanishes along that particular observation direction. This saturates the bound
because we get zero energy along that direction. For that superposition of T and O the
energy along other directions is still nonzero.
A similar argument helps us understand why we also saturate the hT T J i correlator
bound in the four dimensional theory of a Weyl fermion (see section 5). In that case we
can make a superposition of the state T1 in figure 2(b) with the state J1 in 2(e). Notice
that we are using that J couples to a chiral fermion. If there was another fermion with the
same helicity but opposite charge, as it would be the case for a vectorlike current, then we
would have an additional contribution to the state created by the current that will have
a relative minus sign compared to the other charged particle pair. On the other hand, for
the state created by the stress tensor these two contributions have the same sign, therefore
we cannot destructively interfere them.
HJEP1(207)3
This highlights that the bound comes from a quantum mechanical interference effect.
We saturate the bound through a destructive interference effect that prevents particles from
going into a particular direction. It is important to note that this is an interference for the
pair of particles. For example, if we consider a theory of scalars with a U(1) symmetry
generated by a current J , then in a basis of real scalars the current will create two different
scalars, say φ1 and φ2. This cannot interfere with the state created by the stress tensor
where we have the same scalar for the two particles indicated in figure 2(a).
4
Bounds on T T O in d = 3
In this section we will consider the case of d = 3 separately. There are two reasons for doing
this. First, the stresstensor threepoint function has two parity even structures, instead
of three as in d ≥ 4, and has a parity odd piece which is special to d = 3. Secondly, the
correlation function hT T Oi also has an extra parity odd structure special to d = 3 [20].
First we consider external states created by the stresstensor. We parametrize the
threepoint function of energymomentum tensors as
HJEP1(207)3
To obtain a bound on these parameters we can consider a state created by Ψi =
v1T (q, λ0)i + v2T (q, λ1)i. The energy matrix becomes
where T1 = 1 − t4/4, T0 = 1 + t4/4 and Todd = d4/4. These parameters were computed
in [8] in terms of the hT T T i parameters NB, NF and Nodd obtaining
CT T1 =
3
16π2 NF ,
CT T0 =
3
16π2 NB ,
CT Todd =
3
16π2 Nodd .
8We identify our Nodd with their π4pT /3.
hT T T i = NBhT T T iB + NF hT T T iF + NoddhT T T iodd ,
where NB and NF already appeared in the d ≥ 4 case and Nodd parametrizes a new
structure. We use the same convention for the explicit expression for hT T T iodd as in [8].8
In d ≥ 4 the energy onepoint function of the collider experiment has a SO(d − 2) symmetry
for the calorimeter direction n. The linearly independent tensor polarizations are organized
as scalar, vectors or tensors with respect to this symmetry. In d = 3 the group becomes
SO(1) and there are only two types of polarizations, which we take as
λ0 = √
1
2
The collider energy onepoint function for an arbitrary polarization has the structure
εij (ninmλjmλ∗kpnknp + ninmλj∗mλkpnknp) #
2λ2
(4.1)
(4.2)
.
(4.3)
(4.4)
(4.5)
No2dd ≤ NBNF .
For supersymmetric CFTs t4 = 0 just as in the case d ≥ 4. Also, CFTs dual to Einstein
gravity have t4 = d4 = 0.
The average null energy condition implies that the matrix (4.4) is positive definite.
This implies t4 and d4 lie inside a circle t24 + d24 ≤ 42, or equivalently NB ≥ 0, NF ≥ 0, and
Now we will generalize this construction along the same lines as presented in section 3.
We will consider a superposition between stress tensor and a scalar operator states
Ψi = v1T (q, λ0)i + v2T (q, λ1)i + v3O(q)i .
HJEP1(207)3
(4.6)
(4.7)
(4.8)
(4.9)
(4.10)
(4.11)
(4.12)
As anticipated above, for d = 3 the correlation function hT T Oi is now determined by two
parameters
hT T Oi = CTevTeOnhT T Oieven + CTodTdOhT T Oiodd ,
by the following OPE
where the even part is given by specializing the arbitrary d correlator d = 3, and our
choice of normalization for the odd part is given explicitly in appendix D. We can make
our conventions for this latter term as in (3.5) in the following way. We can define CTodTdO
Tzz(z, z¯, y = 0)Tzy(0) ∼ CTodTdO 4z O(0) ,
z¯
where the three spatial coordinates are (z, z¯, y).
Using this normalization, the energy onepoint function is given in terms of a
threebythree matrix as
√CCTevTTeOnCO∗ he3vden(Δ) √CCTodTTdOCO∗ h3oddd(Δ)
Todd
T1
√CCCTeoTvTdeCdOnO
√CTTTCOO
1
he3vden(Δ)
ho3ddd(Δ) v ,
where the functions he3vden(Δ) and ho3ddd(Δ) can be obtained repeating the procedure
reviewed in appendix B and we obtain
ho3ddd(Δ) =
he3vden(Δ) =
12√6π2pΓ(2Δ − 1)
Γ( Δ2+1 )Γ(Δ + 3)
12√6π2pΓ(2Δ − 1)
1
Γ( 7−2Δ ) ,
1
Γ(2 + Δ2 )Γ(Δ + 3) Γ(3 − 2
Δ ) .
Demanding positive definiteness of the energy matrix gives several types of constraints
which involve the scalar OPE coefficients. Two of these bounds are easy to generalize to
an arbitrary number of scalar operators
i
COi
X CTevTeOni 2 feven(Δi) ≤ NB ,
i
COi
X CTodTdOi 2 fodd(Δi) ≤ NF ,
determinant of the 3 × 3 matrix. This gives an independent bound which together with
the bound on hT T T i is sufficient for the positivity of the energy onepoint function
NB
CTevTeOni 2feven(Δi)
CT COi
+ NF
CTodTdOi 2fodd(Δi)
CT COi
−2Nodd
Re [CTevTeOni pfeven(Δi)CTodTdOi
pfodd(Δi)]
CT COi
≤ NBNF − No2dd .
(4.13)
This bound can also be generalized to include an arbitrary number of scalar operators.
However, as opposed to the situation in section 3, the bounds involving different number of
operators are independent. Their expressions in this case become more cumbersome and
we will omit them here.
The (4.10) (4.11) have similar properties as the one appearing for the d ≥ 4 bound.
Namely they diverge at the unitarity bound Δ = 1/2 and have zeros at 6 + 2n (even) and
7 + 2n (odd) for integer n. The zeros in the even case were explained by the existence of
operators with two stress tensors in theories that are dual to weakly coupled gravity, see
the last point in section 3.1. The odd ones have the same explanation, except that now
the scalar operators have the structure ǫABC TAD∂2n∂C TBD.
4.1
ChernSimons matter theories
In this section we apply the bounds derived to large N Chern Simons theories at level k
coupled to fundamental matter. For definiteness we will consider fundamental fermions.
We will denote the ’t Hooft coupling by θ = πN/2k. The elements of the energy matrix
involving the stress tensor were computed in [8] using the explicit large N expressions for
the stress tensor threepoint function [21]. The result is
T1 = 2 cos2 θ ,
T0 = 2 sin2 θ ,
Todd = 2 sin θ cos θ .
(4.14)
Using the conventions in, for example, [22] we can compute the offdiagonal elements
involving stresstensor mixed with a scalar operator. In the fermionic theory we consider the
scalar denoted by O ∼ ψψ¯ has dimension Δ = 2. The final result for the energy matrix is
As a function of the ’t Hooft coupling, this matrix has the property that all the minors
have vanishing determinant.
This implies saturation for all types of superposition of
states. For the case of the stress tensor this was noted in [8], but we find that this is a
more general feature for states where we also act with O.
Even though we do not have a concrete physical picture explaining this, we expect a
picture along the lines of section 3.2, where the interaction with the ChernSimons gauge
field has the effect of replacing free bosons or fermions by “free anyons”.
TT −5
C
−10
−15
−15
−10
−5
0
5
10
15
CT T ε/Cfree
(4.16)
(4.17)
(4.18)
HJEP1(207)3
This discussion can also be applied to the case of CS coupled to fundamental bosons.
From [21] we know that the energy matrix, given in terms of CFT threepoint functions,
can be obtained from the fermionic theory by the replacement θ → θ + π2 when we consider
the operator O ∼ φ2 of dimension Δ = 1. More generally we can consider the answer (4.15)
as giving the energy matrix of a large N theory with a slightly broken higher spin symmetry
parametrized by θ.
4.2
3d Ising model
As another example, we can apply our bounds to obtain predictions for threepoint
coefficients for the 3d Ising model. First let us parameterize the threepoint coefficients of
the energymomentum tensor. Since this theory is parity preserving the coefficient Nodd
computed numerically using the conformal bootstrap in [9, 23]. Explicitly9
in (4.1) is necessarily zero. The remaining two structures in hT T T i have recently been
NB ≈ .9334 ,
NF ≈ .0131 .
The Ising model has a Z2 global symmetry under which T is even. Therefore only Z
2
even scalars participate in the bound. The lightest Z2 even and parity even scalar is the
operator ε whose dimension is known
Therefore, in a normalization where the twopoint function coefficient of ε is one, we can
evaluate (4.12) and find the bound
CT T ε ≤ .0088 = (1.751) Cfree ,
9In making these estimates we use a value of θ ≈ .014. This is the central value of the calculation of [9]
based on expectations for the parity odd scalar gap.
where in the last equation we normalized the answer by the expression (3.15) for the value
of the OPE coefficient in the free theory Cfree = CT T :φ2:/pC:φ2: . Note that although
: φ2 : saturates the bound in the free field theory, the dimension of ε is larger than that of
: φ2 : and hence the OPE coefficient CT T ε may be larger than CT T :φ2:.
We can obtain a stronger bound by including the operator ε′ of dimension Δε′ ≈ 3.8303
in the sum of (4.12).
Using the correct values for feven(Δ) for these dimensions and
normalizing by the T T : φ2 : OPE we obtain the constraint
0.3267CT T ε2 + 0.0063CT T ε′ 2 ≤ Cf2ree .
(4.19)
HJEP1(207)3
Since the operators ε and ε′ are hermitian their OPE coefficients are real and the bound
above defines the allowed region of OPE coefficients as the interior of an ellipse shown in
figure 3.
5
Bounds on T T J in d = 4
As a final example, we consider states created by a linear combination of the
energymomentum tensor and a conserved vector current J in d = 4 spacetime dimensions. In this
case the threepoint function hT T J i is controlled by a single OPE coefficient CT T J and is
parity violating. This threepoint function is presented in detail in appendix E.
One reason why this OPE coefficient is interesting is that it is equivalent to a nontrivial
mixed anomaly between the flavor symmetry generated by J and the Lorentz symmetry
generated by T [24]. In the presence of a background metric g, the current J is not
conserved but instead obeys [
25–27
]
h∇µ Jµ i[g] =
768π2
CT T J ǫµνρσ Rµνδγ Rρσδγ ,
where Rµνρσ is the Reimann tensor.
the net chirality of the charges of elementary Weyl fermions:
In the above, our normalization is such that the coefficient CT T J may be expressed as
CT T J =
X
qi −
X
qj .
Left Weyl i
Right Weyl j
In particular, for the theory of a single Weyl fermion CT T J is one. In an abstract CFT
without a Lagrangian presentation our normalization of the OPE coefficient is defined as follows.
Fix complex coordinates (z, w). Then the OPE of operators restricted to the w = 0 plane is
We will also need the threepoint function hT J J i. This correlator is controlled by two
independent coefficients:
CT T J
Tww(z)Tw¯w¯(0) ∼ 4π6z6 zJ z¯ − z¯J z .
(5.1)
(5.2)
(5.3)
(5.4)
Here the structures CB and W F are those found for the U(1) current in a theory of
free complex bosons (CB) or free Weyl fermions (W F ). In a free field theory, these are
expressed in terms of the charges of elementary fields as (see [15])
2
QCB =
X
complex scalars i
qi2 ,
2
QW F =
X
Weyl fermions i
qi2 .
In general, a single linear combination of these OPE coefficients is fixed by the Ward
identity. We have
hJ J i ∝ CJ ≡ 3
1
QCB + 2Q2W F .
2
The twopoint function coefficient CJ can also be interpreted as a conformal anomaly.
Indeed, in the presence of a nontrivial background gauge field that couples to J , the
energymomentum tensor acquires an anomalous trace. In our conventions this is
By repeating the collider calculation we find that the new offdiagonal matrix element
is given by
hT (q, λT )E (n)J (q, λJ )i =
q
4π
π4 √
CT CJ
r 5
CT T J εijkλ∗T,imλJ,knmnj
!
.
Note that this structure is parity odd as expected. There are other allowed parity odd
expressions in terms of λij and ni, but they do not arise in the nullenergy matrix element.
An important feature of (5.10) is that only those states of SO(2) charge ±1 can mix with
the energymomentum tensor. In particular, this means that bound will only involve the
coefficient T1 defined in (2.10).
Explicitly choosing appropriate polarization tensors we then find that positivity of the
null energy matrix E leads to a single constraint on these OPE coefficients:
CT2 T J ≤ QW F NW F ,
2
hTµ µ i[A] =
4
CJ F αβFαβ .
We can bound the anomaly coefficient CT T J using the same methods described in
earlier sections for scalar operators. We enforce positivity of the average null energy operator
E in the state Ψi created by a linear combination of T and J
Ψi = T (q, λT )i + J (q, λJ )i.
The expectation values hE iλT ·T and hE iλJ ·J have been computed in [1]. The matrix of
energy expectation values in the states Ψi may again be decomposed in terms of the
SO(2) rotation symmetry about the vector n. The current operator J contributes states of
charge −1, 0, 1. As in the review of section 2.2.1 we may express the null energy expectation
value as (qJi/4π) where i is the SO(2) charge. One then finds
J±1 =
Q2W F .
CJ
(5.5)
(5.6)
(5.7)
(5.8)
(5.9)
(5.10)
(5.11)
where NW F = NF /2 counts the effective number of Weyl Fermions in the hT T T i correlation
function. This bound is saturated in the free field theory of Weyl fermions. This can be
understood using the interference argument described in section 3.2.
5.1
Supersymmetry and the Rcurrent
As in our analysis of scalar operators, we can generalize these results to states created by
multiple currents. This is particularly interesting in the case of supersymmetric theories.
In supersymmetric theories, there is always a current JR contained in the same
supermultiplet as T . In particular, since it resides in a different multiplet it can be distinguished
from an ordinary flavor current JF . We would like to improve our bound on the trace
anomaly of JF to account for the fact that the Rcurrent JR always exists. In order to do
this we consider the state created by
Ψi = v1T (q, λT )i + v2JR(q, λJ )i + v3JF (q, λJ )i .
(5.12)
The new ingredient appearing in the calculation of the energy matrix corresponding to
this state involves the threepoint function hT JRJF i. Using superconformal invariance we
can fix this correlator completely. Since the details are not very illuminating we will
outline the procedure.
The number of parity even structures, two of them, coincides
with the ones appearing in hT J J i, namely relaxing permutation symmetry does not add
new structures [17]. Moreover, using supersymmetric Ward identities [28] one can check
that no parity odd structure is allowed for hT JRJF i.10 Out of the two OPE coefficients
characterizing hT JRJF i, a linear combination of them is related to the twopoint function
hJRJF i, which vanishes due to superconformal invariance. This leaves hT JRJF i fixed by a
single OPE coefficient. Finally, since JR lies in the same multiplet as the stress tensor we
can relate this number to CT T F , the mixed anomaly generated by the flavor current.
Combining the results outlined in the previous paragraph, and the fact that there is
no new structure involved in the collider calculation, it is straightforward to obtain the
offdiagonal matrix element
hJR(q, λJ )E JF (q, λJ )i =
q
4π
r 20
3π4 √
CT T F
CT CF
!
,
where we chose n = (1, 0, 0) and λJ = (0, 1, i) for definiteness.
We can express parameters related to the Rcurrent in terms of a and c = CT π4/40.
energy matrix as a function only of a, c, CT T F and CF . We obtain
The twopoint function is related to CT by a supersymmetry Ward identity as CR = 136 c.
Its mixed anomaly is also fixed by supersymmetry to CT T R = 16(c − a). Finally the energy
onepoint function is given by J±R1 = ac [1, 7]. Supersymmetry also fixes this parameter for
flavor currents as J±F1 = 1. Taking these facts into account allows us to write down the
(5.13)
(5.14)
hΨE Ψi =
q
4π
v†
2c−a
c
√3 c−c a
√3 c−c a
a
c
√12c C√TCTFF √16c C√TCTFF
√12c C√TCTFF
√6c C√TCTFF v ,
1
1
SO(2) spin.
in a generic theory.
where for definiteness we have chosen λJ = (0, 1, i) and a tensor polarization with the same
10This is not true for a threepoint function of a stress tensor and two different conserved currents hT J1J2i
Enforcing the positivity of this matrix yields several constraints. The leading
twobytwo minor involving states T (q, λT )i + JR(q, λJ )i gives the bound
1
a
2 ≤ c ≤ 2
3
,
which coincides with those derived in [1]. This bound is saturated by a free chiral multiplet,
ac = 21 , or a free vector multiplet, ac = 32
To constrain the gravitational anomaly coefficient we evaluate the determinant of the
full threebythree matrix (5.14). This gives the following bound on the mixed anomaly for
a flavor current
a
c − 2
1
36c − 24a −
CT2 T F
CF
≥ 0 .
For a free chiral multiplet the bound is automatically saturated, since the first term in
the left hand side vanishes independently of CT T F . Therefore we will assume that ac > 21 .
Then we obtain the following bound
CT2 T F
12 CF
≤ 3c − 2a ,
(5.15)
(5.16)
(5.17)
which is stronger than the one derived in the previous section, without the use of
supersymmetry. Note also that this is consistent with the free vector multiplet. In that case the
righthandside vanishes, but there are also no flavor currents.
To conclude this section, we can mention some contexts where such bound on the
mixed anomaly is relevant. First of all, when we consider holographic CFT this anomaly
is related to a 5d ChernSimons term of the form R A ∧ R ∧ R, where A is the gauge field
dual to the current J (we will see in the next section how our bounds translate to bounds
on the gravity couplings for the case of T T O).
Finally, in the context of hydrodynamics and transport, quantum anomalies induce a
special type of transport coefficients, see [29] and, in particular, for the mixed anomaly [30–
32]. The coefficient bounded in this section CT T J , is related to the mixed anomaly recently
observed experimentally in Weyl semimetals [33]. In the linear response regime, the mixed
anomaly produces an energy current ~j given by [30–32]
~j = 24CT T J T 2B~ ,
(5.18)
where we denote the temperature by T and the system is placed in a fixed magnetic field
B~ . This allows us to translate our results into concrete bounds for transport coefficients.
6
Bounds on coefficients of the AdS effective action
If the d dimensional boundary theory has an AdSd+1 dual, then we would like to translate
the bounds on CT T O to bounds on the coefficients of the bulk effective action. We are
imagining that the theory has a large N expansion. Then, to leading order, the bulk is
given by a collection of free fields propagating on the AdS metric. The simplest interactions
correspond to bulk threepoint interactions. These lead to threepoint functions in the
boundary theory. For the case of gravitons we have a threepoint interaction coming from
the Einstein Lagrangian, but it is also necessary to include higher derivative terms, of the
form W 2 and W 3, in order to get the most general structures for the tensor threepoint
function. It is possible to match the coefficients of the new structures to the coefficients of
these higher derivative terms in the Lagrangian [
1, 3
].
Here we consider the same problem for the case of the hT T Oi correlator. The first
observation is that in Einstein gravity this correlator is zero, since the action of any field,
expanded around the minimum of its potential has an action without a linear term in the
scalar field. Notice that this also implies that a massive scalar field cannot not decay into
two gravitons. However, we can write the higher derivative term
S = Mpdl−1α Z
dd+1x√gχW 2
in the action, where we normalized the χ field to be dimensionless.11 This term enables the
field χ to decay into two gravitons. In flat space there is only one structure for the on shell
threepoint function between a scalar and two gravitons, except in four dimensions where
threre is also a parity odd one, as we discuss later. Therefore the vertex (6.1) represents
the general interaction that we can have in the theory. There can be other ways to write
it which give the same threepoint function as (6.1). It is possible to check that (6.1) gives
rise to a hT T Oi threepoint function with the coefficient
CT T O
pf (Δ)
√
CT
d + 1 Γ(d/2) RA2dS
α
.
(6.1)
(6.2)
At first sight, it seems surprising that the function f (Δ) appearing here is the same as the
one that appears in the bound (3.10). This means that the Δ dependence disappears when
we express the bound in terms of α. This is easy to understand when we derive (6.2) as
follows.
First we notice that integrating the stress tensor along a null line, as in the definition
of the energy measurement E = R dx−T−−(x−, x+ = 0, ~y = 0), we produce a shock wave
in the bulk that is localized at x+ = 0. We can then imagine scattering a superposition
of χ and a graviton through this shock wave. This leads to a time delay that is given by
a matrix mixing the graviton and the scalar. An important point is that the propagation
through the shock wave is given by integrating the wave equation in a small interval before
and after x+ = 0. Only the shock wave contributes to this short integral over x+, but
the scalar mass term does not contribute. Therefore the time delay matrix is independent
of the mass of the scalar. We can determine the precise coefficient in (6.2) by doing this
explicit computation for Einstein gravity plus (6.1). We then get a bound on α by requiring
that the time delay is positive. Comparing this to the bound (3.10) we fix the coefficient
to the one in (6.2). We explain this in more detail in appendix F.
This same shock wave method enables one to set even stricter bounds on α if one
assumes that there is a gap to the higher spin particles,12 as was discussed in [34] for the
Mpdl−1
2
R dd+1x√gR. Similarly, the action of the scalar field is S = Mpdl−1
2
12We thank E. Perlmutter and D. Meltzer for discussions on this issue.
R [(∇χ)2 − m2χ2].
11Here Mpl is the reduced Planck mass in d + 1 dimensions, defined so that the Einstein term is S =
case of the graviton higher derivative interactions. A similar analysis can be done for the
5d ChernSimons term coupling dual to the mixed anomaly [35].
In string theory, we expect that α is the order of α′, the inverse string tension. If
gravity is a good approximation, α′ ≪ R2, then we find that the bound on (6.2) is far from
being saturated. The bound is saturated only as the string length becomes of the order of
the radius of AdS. In particular, this implies that the bound is satisfied, and far from being
saturated, for the Konishi operator of N = 4 super Yang Mills at strong coupling. This
operator is the lightest nonprotected single trace operator which has a dimension growing
like Δ ∝ λ1/4 at strong coupling, λ ≫ 1.
In the four dimensional case, we can also have a parity odd correlator with a
corresponding coupling. In flat space this is related to the fact that the threepoint functions with
++ or −− graviton helicities are Lorentz invariant by themselves. (The −+ graviton
helicities are forbidden by angular momentum conservation). We can then write the action as
S = Mp2l Z
(R − 2Λ) + [(∇χ)2 − m2χ2] +
αeχW 2 + αoχW W ∗ ,
(6.3)
where as above we have defined χ to be dimensionless.13 In this normalization αi has
dimensions of length squared. They can be related to the coefficients of the threepoint function as
(6.4)
(6.5)
CTevTeOnheven(Δ)
√
CT
= √
24
αe
,
CTodTdOhodd(Δ)
√
CT
= √
24
αo
The bounds in this case then read
pαe2 + α2
o
RA2dS4
1
≤ 12√2 ,
in the case that there are no purely gravitational corrections to Einstein gravity. Of
course, if there are threepoint functions that lead to corrections to Einstein gravity, then
the bound is corrected to those given in section 4.
7
Constraints for deSitter and inflation
The physics of inflation might be our very best window into very high energy physics.
The standard inflationary theory starts with a scalar field coupled to the Einstein action
and includes all two (or less) derivative interactions. The universe undergoes a period of
expansion that is governed by a nearly deSitter solution, characterized by a Hubble scale H
that is nearly constant. The effective coupling of the gravitational sector is of order H/Mpl
which is very small, less than 10−5. However, it is possible that there are corrections to
the two derivative action due to the presence of a light string scale. The value of the string
tension could be fairly low H2 . T . When the string tension becomes comparable to the
Hubble scale, we expect significant corrections to the two derivative action. We do not have
an explicit scenario where this happens. However, a similar situation happens in AdS space
13We also define (W ∗)μνρσ = 21 ǫμνδγW δγρσ.
when we consider a gravity dual of a not so strongly coupled large N theory. Therefore it is
natural to question whether something similar could happen in inflation and we can look for
signatures of such a low string scale. It is important to find signatures that are as model
independent as possible. Specially nice signatures are those that have a nonvanishing
contribution in the deSitter approximation. These are not strongly suppressed by slow
roll factors. In addition, their form is strongly constrained by the deSitter isometries.
An example of such contributions are the threepoint functions of gravity fluctuations,
where the higher derivative corrections were discussed in [36]. Another interesting case are
the couplings of the form fe(χ)W 2 or f0(χ)W W ∗. These two couplings are particularly
interesting because their effects are visible at the twopoint function level.
Let us discuss first the parity odd coupling, which leads to chiral gravity waves [37, 38].
Namely, we have different gravity wave twopoint functions, hhL, hhR, for the left and right
handed circularly polarization. We can define the asymmetry A as
A ≡ hhL + hhR = 4π f˙o(χ) H2 = ±4π√2ǫ
hhL − hhR
H
∂f
∂χ
H2 ,
χ =
φ
Mpl
,
(7.1)
where χ is defined to be dimensionless and φ is the inflaton with canonical normalization.
(The ± comes from going from χ˙ to √ǫ, since the derivative of the scalar can have either
sign). If we were in AdS4 we would have a sharp bound on the coefficients via the
condition (6.5), after we identify αo = ∂∂χf . It is reasonable to think that in the deSitter case too,
there will be trouble is the bound is violated. Of course, we know that even nearsaturation
of the bound implies that the field theory approximation is breaking down.
In the deSitter case we do not have a sharp derivation of a bound from boundary theory
reasoning. We do not have an analog of the null energy condition, discussed in section 4,
for the boundary theory, since the boundary theory is purely spacelike. It would be nice
to have a sharp derivation of a deSitter version of the bound. In deSitter, we can talk
of a “quasibound”, which we get by simply applying the same bound on the coefficients
of the action that we had in antideSitter. This quasibound should be viewed simply
as an educated guess, including numerical coefficients, for the validity of bulk effective
theory. A near saturation of these quasibounds is a strong indication of a light string scale
which could also have other manifestations such as indirect evidence of higher spin massive
particles, etc [39]. In summary, in deSitter also we have a quasibound on the coefficients
similar to (6.5), with 1/RAdS → H
A ≤ 12
4π √ǫ .
= pαe2 + αo2 ≤ 12√2 .
H2
This bound, then implies a quasibound on the asymmetry (7.1) of the form
The allowed values by this quasibound seem to be smaller than the smallest possible
measurable value from the CMB Bmodes as analyzed in [40]. Conversely, this means that
(7.2)
(7.3)
if chiral gravity waves through EB mode correlators are measured, then we would need a
higher derivative coupling with a coefficient so large that it violates (7.2).
Let us turn now to a discussion of the parity even coupling. This coupling gives rise
to a violation of the consistency condition for the twopoint function [41], even in the case
that the speed of sound is close to one,
where we assumed that the speed of sound for the scalar is close to one. Here nt is the tensor
spectral index and r the tensor to scalar ratio conventionally defined. Then the bound we
had in (7.2) translates into the following constraint on the violation of the consistency
HJEP1(207)3
condition
−8 nt
Comments on scalartensortensor threepoint functions
The φW 2 higher derivative coupling between the scalar and the graviton also give rise
to new contributions to the scalartensortensor threepoint function. This is a
contribution, that is nonvanishing in the deSitter limit. More precisely, if we can approximate
∂χfe(χ(t)) by a constant, then we get a contribution even in deSitter space. The standard
Einstein gravity contribution, [42], is suppressed by a slow roll factor √ǫ, if we assume that
∂χf is of order one. Of course, our bound constrains the size of this threepoint function
because it is constraining the size of the coefficient αe ∼ ∂χfe(χ(t)).
The threepoint function for the parity odd coupling fo(χ)W W ∗ was computed in [43],
where it was found to be proportional to ∂χ2f . One might have naively expected a deSitter
invariant contribution proportional to α0 = ∂χf0, when we approximate this by a constant.
The explicit computation by [43] shows that there is no such contribution. This seems
surprising at first sight because this parity odd coupling does indeed give a nonvanishing
contribution to the threepoint function in the AdS4 case. The reason it vanishes in
deSitter is that it gives a contribution to the deSitter wavefunction that is a pure phase,
which disappears when we take the absolute value squared of the wavefunction. The same
happens with the W 2W ∗ parity violating graviton threepoint coupling [44]. The correlator
proportional to ∂χ2f found in [43] has an extra factor of φ˙ and is not expected to be deSitter
invariant (though we did not check this explicitly).
It should be noted that the correction to the twopoint function consistency
condition (7.4) has the right form so that the consistency condition involving the soft limit of
the threepoint function [42, 45] is obeyed, though we have not explicitly checked the
precise numerical coefficients. A similar remark applies in the parity odd case; the correction
to the twopoint function (7.1) is such that the soft limit of the threepoint function in [43]
obeys the consistency condition.
Acknowledgments
We thank H. Casini, S. Giombi, R. Meyer, E. Perlmutter, D. SimmonsDuffin, and D.
Stanford for discussions. We also thank E. Perlmutter for comments on a draft. C.C. is
supported by the Marvin L. Goldberger Membership at the Institute for Advanced Study,
and DOE grant DESC0009988. J.M. is supported in part by U.S. Department of Energy
grant DESC0009988 and the Simons Foundation grant 385600.
A
Absence of positive local operators
Let us review the essential steps of [11] in a modern language. Let Φ be any Hermitian
operator and 0i the Lorentz invariant vacuum state. We make two assumptions:
• The onepoint function h0Φ0i vanishes.
• For all states ψi in the Hilbert space, the expectation value hψΦψi is nonnegative.
Under these assumptions we may prove that Φ annihilates the vacuum state, Φ0i = 0.
Indeed, for any positive operator, the CauchySchwarz inequality implies that
hψΦ0i2 ≤ hψΦψih0Φ0i .
Since the righthand side is zero by hypothesis, we conclude that Φ0i must vanish.
If we now further assume that Φ is an operator localized within a compact region R,
we can deduce that Φ must vanish. To demonstrate this, consider operators localized in a
region R′ that is spacelike separated from R. Let us denote by OR′ a set (sum of products)
of smeared operators in region R′, then we have
0 = h0OR1′OR2′Φ(z)0i = h0OR1′Φ(z)OR2′0i ,
where we have used that the operator OR2′ is spacelike separated from the region where Φ
is localized in order to move it to the right of Φ. However, according to the ReehSchlieder
theorem [46], any state ψi may be approximated to arbitrary precision by acting with a
(smeared) set of local operators in any open set in spacetime. Since the region of points
that are spacelike separated from a finite compact region is an open set, we may apply this
idea to the righthand side above to conclude that for any sates ψii
hψ2Φψ1i = 0 .
This implies that Φ vanishes as an operator.
the null line.14
It is interesting to pinpoint exactly where this logic breaks down for nonlocal operators
such as the average null energy operator E. As long as the region that is spacelike separated
to Φ is open, one may repeat the ReehSchlieder argument and prove that Φ vanishes even
if it is nonlocal. The way the null energy operator E avoids this conclusion is that it
has support along a complete null line and hence the region of points that are spacelike
separated to E is not open, since it consists just of the codimension one null plane containing
14We that H. Casini for an enlightening discussion on this point.
(A.1)
(A.2)
(A.3)
Details of the collider calculation
In this appendix we will give more details on the calculation of the energy expectation
value for a conformal collider experiment that we considered in this paper, giving a bound
on CT T O in arbitrary dimensions.
B.1
Normalized states
The states that we consider for the collider experiment are superposition of states of
normalized wavepackets. Following [1] we take the state defined as
O(q, λ)i ≡ N
Z
ddx e−iqx0 exp −
x20 + ~x2
σ2
λ · O(x)0i,
q > 0
(B.1)
HJEP1(207)3
(λ0)ij =
r d − 1
d − 2
1
ninj − (d − 1) δij ,
which satisfies Tr(λ0) = 0 and λ0 · λ0 = 1. In this case the normalization condition gives
hT (q, λ0)T (q, λ0)i = 1
NT0
−1/2 = CT
4(d − 1)π d2 +1
Γ d2 Γ(d + 2) 2
q d
,
where the normalization of the twopoint function is
where qσ ≫ 1. We find the normalization N by requiring the state to have unit norm in
this limit. We will give their values only for the operators and polarizations relevant to
computing the bound on CT T O. Namely for a scalar operator and for the stress tensor with
polarization which is scalar with respect to the SO(d − 2) symmetry perpendicular to n.
We normalize the scalar operator such that its twopoint function is hO(x)O(0)i =
COx−2Δ. Then the normalization condition for the state considered in the collider
experiment is
hO(q)O(q)i = 1
NO−2 = C
2π d+22
q 2Δ−d
O Γ(Δ)Γ(Δ − d2 + 1) 2
where we used the following integral identity
Z
ddx
eiqx0
x2Δ =
2π d+22
Γ(Δ)Γ(Δ − d2 + 1) 2
q 2Δ−d ,
q > 0
We will also need the proper normalization for the scalar state created by the stress
tensor, which has the form
T (q, λ0)i ≡ NT0
Z
ddx e−iqx0 (λ0)ij Tij (x)0i,
where we assumed the localized wavepacket limit. If we use conservation of the stress tensor
we can chose the polarization along spatial directions. The normalized scalar polarization is
hTµν (x)Tρσ(0)i =
xC2Td Iµνρσ (x),
(B.2)
(B.3)
(B.4)
(B.5)
(B.6)
(B.7)
and the tensor structure that appears derived in [15] is
Iµν (x) = gµν − 2 xµ xν
Of course, by SO(d − 1) rotational symmetry, the normalizations for T1 and T2 are also
given by (B.6), once the polarizations are normalized to unity. Below we will perform the
experiment of [1] for linear superpositions of these normalized states. But first we will
review the form of the correlators we will need, mainly to fix notation and conventions.
HJEP1(207)3
B.2
Threepoint functions
The threepoint functions we will need are hT T T i, hT OOi and hT T Oi. Their form were
derived in [15] and the first two were studied in the context of the conformal collider in
four dimensions in [1] and generalized to arbitrary dimensions in [
3
]. First, we will focus on
hT T Oi which was not studied previously in the context of the conformal collider. The form
consistent with conformal symmetry and conservation of the stresstensor found in [15] is
CT T O ≡ aˆ + 8(ˆb + cˆ).
hTµν (x1)Tρσ(x2)O(x3)i =
1
x212d−Δx2Δ3x3Δ1 Iµν µ ′ν′ (x13)Iρσ
x223 and the tensor structure is a sum of three terms
tµνρσ (x) ≡ aˆ hµνρσ (x) + ˆb hµν2ρσ (x) + cˆ hµνρσ (x),
1 3
where each hi is traceless and symmetric under µν
↔ ρσ, x → −x
1
hµνρσ (x) =
2
hµνρσ (x) =
1
1
x4 (xµ xν xρxσ + . . .),
x2 (xµ xρgνσ + . . .),
3
hµνρσ (x) = gµρ gνσ + . . . ,
where dots represent terms needed for expressions to be traceless and symmetric and in
each line they involve a fixed number of factors of x. The main advantage of this approach is
that it makes transparent the OPE limit x2 → x1, or equivalently taking the x3 → ∞ limit
1
Tµν (x)Tρσ(0) ∼ x2d−Δ (aˆ hµνρσ (x) + ˆb hµν2ρσ (x) + cˆ hµνρσ (x))O(0).
1 3
Conservation can be imposed in this limit to the right hand side to the equation above,
giving the two independent relations
aˆ + 4ˆb − 2
1
(d − Δ)(d − 1)(aˆ + 4ˆb) − dΔˆb = 0,
aˆ + 4ˆb + d(d − Δ)ˆb + d(2d − Δ)cˆ = 0.
This fixes the three point function to a single conserved structure up to an overall
coefficient, which we can define as
(B.8)
(B.9)
(B.10)
(B.11)
(B.12)
(B.13)
(B.14)
(B.15)
(B.16)
(B.17)
(B.18)
V1 =
V2 =
(Z1 · X2)(X1 · X3) − (Z1 · X3)(X1 · X2)
(Z2 · X3)(X2 · X1) − (Z2 · X1)(X2 · X3)
X2 · X3
X1 · X3
H12 = −2((Z1 · Z2)(X1 · X2) − (Z1 · X2)(Z2 · X1)).
Starting from the expression in d + 2dimensional embedding space we can obtain the
ddimensional correlator hT (x1, z1)T (x2, z2)O(x3)i by using the replacements −2Xi · Xj →
xi2j, Zi · Zj → zi · zj and Xi · Zj → xij · zj, where now in ddimensions we define T (x, z) =
zµ zνTµν (x), with the index running from 0 to d − 1. Of course after these replacements
the answer coincides with the OsbornPetkou threepoint function.
We can match the
coefficients of the different representations by taking the OPE limit. The result gives
Another standard way of representing conformal threepoint functions is given by the
spinning correlator formalism of [17]. We will write the correlator in terms of the embedding
space coordinate Xi ∈ Rd+1,1 and the polarization Zi ∈ Rd+1,1 such that Z2 = 0. The
correlator we need is in appendix A of [17] and in terms of the usual conformal structures
Vi and Hij is given by
hT (X1, Z1)T (X2, Z2)O(X3)i =
α1V12V22 + α3H12V1V2 + α6H122
(−2X1 · X2)d+2− Δ2 (−2X2 · X3) 2 (−2X3 · X1) 2
Δ
Δ , (B.19)
where T (X, Z) ≡ ZAZBTAB(X) (with the index running from 0 to d + 1) and the labels
1, 3, 6 on the α’s correspond to the subset of the 10 structures that hT T Oℓi has for arbitrary
ℓ, that survive in the case ℓ = 0. The building blocks are
2α1 + 2 α3(−4 + d2 − dΔ) + α6dΔ = 0.
The OPE coefficient (B.18) is now CT T O = aˆ + 8(ˆb + cˆ) = α1. The conservation equations
in terms of these parameters are
α1(2 + Δ − d(1 − d + Δ)) + α3 − 2 − Δ + (Δ + 2)
= 0
d
2
α1 = aˆ + 8(ˆb + cˆ),
α3 = 4(ˆb + 2cˆ),
α6 = 2cˆ,
1
1
which we can solve in terms of CT T O.
function. Another correlator we need is
We defined in the main text the notation we will use for the stresstensor threepoint
hTµν (x1)O(x2)O(x3)i = CT OO x1d2x223Δ−dx3d1 Iµνρσ (x13)
X1ρ2X1σ2
X122
− d
1 gρσ ,
dΔ
which is fixed by a Ward identity to be CT OO = −CO (d−1)Ωd−1 , with Ωd being the are of
(B.20)
(B.21)
(B.22)
(B.23)
(B.24)
(B.25)
(B.26)
(B.27)
(B.28)
As explained in the main text we want to consider states of the form
Ψi = v1T (q, λ0)i + v2O(q)i,
where we take v = (v1, v2) ∈ C2 such that v12 + v22 = 1. The energy onepoint function
in the collider experiment is
In this section we will compute the entries of this matrix. The diagonal elements were
already computed in [
1, 3
] and are given by
HJEP1(207)3
hT (q, λ0)E (n)T (q, λ0)i =
q
Ωd−2
d − 3
1 − d − 1 t2 −
d2 − 1 ((d − 3)NF − 2(d − 2)NV )
(d − 3)((d − 1)(dNV + NF ) + 2NB)
,
t4 =
d2 − 1 (NB − NF + NV )
2NB + (d − 1)(dNV + NF )
. (B.32)
Using these expressions we can find the parameters we called T0, T1 and T2 in the main
text. They are given by equation (2.10) where the functions ρi(d) are given by
The state created by a scalar operator gives
ρ0(d) =
ρ1(d) =
ρ2(d) =
1
1
Ω2d−1
Ω2d−1
1
Ω2d−1
d(d + 1)(d − 2)
2(d − 1)
d(d + 1)
4
d(d + 1)(d − 2)
2(d − 3)
q
Ωd−2
.
x2 · m¯ → ∞,
m¯ = (−1, n).
Now we will obtain the offdiagonal element of this matrix we has not been computed
in the literature. To perform the calculation in arbitrary dimensions it is convenient to
use the spinning correlator formalism. Since we are computing an expectation value the
correlator we need to consider is not timeordered. The right iǫ prescription for this purpose
was explained in [1] and [
3
], and we will omit it here to ease the notation. We start from
the ddimensional expression
hT (x1, z1)T (x2, z2)O(x3)i =
α1V12V22 + α3H12V1V2 + α6H122
(x212)d+2− Δ2 (x223) 2 (x231) 2
Δ Δ
We will chose T (x2, z2) to be the insertion taken to infinity and giving E (n). First we take
z2 = m = (1, n) (we chose the mostly plus convention for the metric in Minkowski space).
Therefore T (x2, z2) → 41 T−−(x2). Then we take the limit
(B.29)
(B.30)
(B.31)
(B.33)
(B.34)
(B.35)
(B.36)
(B.37)
(B.38)
To take this limit we can use the results of appendix F of [14], and obtain
lim
hT (x1, z1)rd−2T−−(x2)O(x3)i =
α1Vˆ12Vˆ22 + α3Hˆ12Vˆ1Vˆ2 + α6Hˆ122
2d(x12 · m)d+2− 2 (x21) Δ2 (x2 · m) Δ2
Δ
where the structures in this limit are
Vˆ1 =
x2 · m
x
1
, Vˆ2 =
x1 ·2m , H12 = −z1 · m.
We can get the correct polarization of the insertion T (x1, z1) by replacing zµ zν → λµνT ,
assuming λ is already traceless and symmetric which is true for expression (B.5). To
simplify the expressions we will choose n = (1, 0, . . . , 0) and write the positions as x =
(x+, x−, x⊥), where x+ = x · m¯ = x1 + x0, x− = x · m = x1 − x0 and x⊥ corresponds to
the d − 2 transversal components. Then we can define
x2 →∞
Gˆ = lim (x2+/2)d−2hλ0 · T (x1)T−−(x2+, x2−, 0)O(x3 = 0)i
+
which is given by
4
Gˆ = α1λ1T1 (x1−)2( x41 − x11x1−2x12 + x11x11(x1−2)2 − d −12 Pi⊥(xi1⊥)2(x1−2)2)
Δ
2d(x1−2)d+2− 2 (x21) Δ2 +2(x2−) Δ2 +2
+α3λ1T1
2−d(x1−)(x11x1−2 − 2 1
1 x2)
(x1−2)d+2− Δ2 (x21) Δ2 +1(x2−) Δ2 +1
+ α6λ1T1
2−d
(x1−2)d+2− Δ2 (x21) Δ2 (x2−) Δ2
First we do integral over x2−, using the following identity
1
Z dx2− (x2− − iǫ)b(x1−2 − iǫ)a =
(x1− − 2iǫ)a+b−1 Γ(a)Γ(b)
2πi
Γ(a + b − 1)
,
where we made explicit the pole prescription. Finally, to take the limit of the localized
wavepackets is equivalent to setting x3 → 0 and make a Fourier transform with respect to
x1 with momentum (q, 0, . . . , 0), namely
Z ddx1e−iqx10 Z dx2−Gˆ.
depends on x1+, x1− and x1⊥ · x1⊥. Then the integral can be written as
To do this we first integrate over the d − 2 transverse directions x1⊥ and then integrate
over the lightcone coordinates x±. Because of SO(d − 2) invariance, the integrand only
1
Z ddx1e−iqx10 Z dx2−Gˆ =
Ωd−3 Z dx1+e−i q2 x1+ Z dx1−ei q2 x1−,
(B.39)
(B.40)
(B.41)
(B.42)
(B.43)
(B.44)
HJEP1(207)3
×
Rd−3dR F (x1+, x1−, (x1⊥)2 = R2),
(B.45)
where we defined F = R dx2−Gˆ to indicate the functional dependence explicitly. After
performing these integrals we use conservation conditions to write α3 and α6 in terms of
α1 = CT T O using equations (B.16) and (B.23). Combining the three structures gives
Z
ddx1e−iqx10 Z
dx2−Gˆ =
CT T O22−d(d − 1)π d2 +2Γ(d + 1)
2
(d − 2)Γ Δ2 + 2
q Δ+1
2
(B.46)
In this expression we generalized the answer to arbitrary n by replacing λ11 → λij ninj .
Finally, we need to replace the specific value of the polarization tensor (B.5) and the proper
normalization of the collider states (B.6) and (B.2). The final answer for the offdiagonal
entry of the energy matrix is
HJEP1(207)3
where
q
Ωd−2
CT T O h(Δ),
CT CO
h(Δ) ≡
π 2 Γ(d + 1)qΓ d2 − 1 Γ(d + 2)Γ(Δ)Γ Δ − 2
d+1
d−2
2dΓ d−1 Γ Δ2 + 2
2
2
Γ d+Δ Γ d − 2
Δ
2
Then the energy matrix that gives the expectation value for these superposition states is
Having computed the energy matrix the next step is to impose ANEC, which is
equivalent to imposing positivity of the energy expectation value for the collider experiment.
This means that for all states
Ψ(v)i = v1T (q, λ0)i + v2O(q)i
hΨ(v)E (n)Ψ(v)i > 0,
∀v ∈ C2.
we need to impose
This constraint is equivalent to the positivity of all the leading principal minors of the
energy matrix. The first constraint is
T0 = 1 − d − 1 t2 −
d − 3
which is the same as the original constraints of [1] and [
3
]. The next minor imposes the
positivity of the 2× 2 matrix which is
We can write T0 and CT in terms of the hT T T i structures NB, NF and NV . This gives
the equivalent expression that we quoted in the introduction
(B.47)
(B.48)
(B.49)
(B.50)
(B.51)
(B.52)
(B.53)
(B.54)
(B.55)
where
(d − 1)3dπ2dΓ d2 Γ(d + 1)Γ(Δ)Γ Δ − d −22
Δ 2
These two conditions (B.53) and (B.54) are necessary and sufficient for the energy to be
positive for any state of the form (B.51). For operators O that are not hermitian this
bound does not have information about the phase of the OPE coefficient CT T O.
C
Free scalar correlators
In this section we will present some details on the calculation of the T T O correlators for a
free scalar that saturates the bound above. We use the normalization of [15] for
(B.56)
(C.1)
(C.2)
(C.3)
(C.4)
(C.5)
(C.6)
1
2
and the stress tensor is defined as
hφ(x)φ(0)i =
(d − 2)Ωd−1 xd−2
1
1
.
Tµν = :∂µ φ∂νφ: − 4(d − 1)
((d − 2)∂µ ∂ν + gµν ∂2) :φ2: .
Since scalar operators have integer dimensions we only need to consider O such that Δ < 2d.
The first one is O ∼ φ. This one is predicted to vanish since f (Δ = d −22 ) → ∞. This
is indeed the case since an odd number of fields appear in hT T φi. The next operator is
O = :φ2: of dimension Δ = d−2. The correct normalization of the twopoint function gives
h :φ2: (x) :φ2: (0)i =
(d − 2)2Ω2d−1 x2(d−2)
1
, C
O =
2
(d − 2)2Ω2d−1
.
Using Wick contractions we can also compute hT T φ2i. One can check that the answer
has the conformal invariant structure (B.19) with
CT T O = α1 =
(d − 2)d2 1
2(d − 1)2 Ω3d−1
, α3 = − d − 2 α1, α6 =
(d − 2)d α1
4
2
Finally, the function appearing in the bound takes the value
Putting everything together we find that the bound is saturated
f (Δ = d − 2) =
8(d − 1)4π2d
(d − 2)4Γ d2 + 1
4
C
O
2
CT T O f (d − 2) = 1 ≤ NB = 1.
We have seen that the bound is saturated by a scalar field with O = :φ2:. Nevertheless
there is one more primary scalar field we can make of dimension less than 2d, namely
O = : φ4 : which has dimension Δφ4 = 2(d − 2). Working out the Wick contractions we
can verify that hT T : φ4: i = 0. We can argue more generally that this is so. The form of
the correlator (B.19) indicates that there is a x1 → x2 singularity whenever Δ < 2d. On
the other hand, to have a nonzero answer we can only take a Wick contraction which is
between T and :φ4: but not between the stress tensors. Therefore if this calculation would
give a nonzero answer, it will be finite when x1 → x2. The only way this is consistent with
the form of the correlator fixed by conformal symmetry (B.19) is if it indeed vanishes.
In dimensions d ≥ 4 the threepoint function hT T Oi has only a parityeven structure
consistent with permutation symmetry and conservation of the stresstensors. The situation
for d = 3 is special since only for this number of dimensions a new parityodd structure
appears, that is also consistent with all the requirements. In this case the full correlator is
hT T Oi = hT T Oieven + hT T Oiodd
where the parityeven part coincides with the answer for d ≥ 4
(D.1)
(D.2)
(D.3)
(D.4)
(D.5)
(D.6)
hT (X1, Z1)T (X2, Z2)O(X3)ieven =
and the new structure is
hT (X1, Z1)T (X2, Z2)O(X3)iodd =
α1V12V22 + α3H12V1V2 + α6H122 ,
Δ Δ Δ
β1V1V2 + β2HΔ12 ǫ(Z1, Z2, X1, X2, X3),
Δ Δ
Since conservation put constraints independently for α1,3,6 and β1,2 we can forget about
the parityeven part and we get β1(Δ − 3) − β2(Δ + 1) = 0. Therefore the parityodd
structure is also fixed by a single OPE coefficient which we denote CTodTdO, as opposed to
the one in the even part CTevTeOn = α1. For completeness we present the same correlator in
the Osborn and Petkou formalism
hTµν (x1)Tρσ(x2)O(x3)i =
1
x162−Δx2Δ3x3Δ1 Iµν µ ′ν′ (x13)Iρσ
1
tµνρσ (X) =
tµνρσ (X) = δµρ xγ
2
xµ xρxγ
x3
x
ενσγ + . . . ,
ενσγ + . . . ,
where the dots represent other terms of the same form to make the answer symmetric,
traceless and permutation symmetric. Conservation imposes eˆ(Δ − 7) + dˆ(Δ − 3) = 0.
Either in the Osborn and Petkou formalism or in the spinning correlator formalism, we
define the parityodd OPE coefficient as CTodTdO ≡ dˆ+ eˆ = (β2 − β1)/4.
E
hT T J i threepoint function
In this appendix we will provide some details on the CFT threepoint function controlling
the mixed gaugegravitational anomaly hT T J i. Imposing permutation symmetry between
the stresstensors and conservation, this correlator only involves an allowed parityodd
structure. In the spinning correlator formalism it is given by
hT (X1, Z1)T (X2, Z2)J (X3, Z3)i ∼
X142X223X321
H12 − 4V1V2 ǫ(Z1, Z2, Z3, X1, X2, X3),
(E.1)
Petkou. This can be written as
hTµν (x1)Tρσ(x2)Jα(x3)i =
x23
explicitly
and
where as usual the uppercase denote coordinates in embedding space and Hij and Vi
are the usual structures defined in [17]. From this expression it is possible to deduce the
conservation equation for the current when the CFT is placed on a curved background and
gives the right normalization for CT T J [24]. Then the threepoint function is
hT (X1, Z1)T (X2, Z2)J (X3, Z3)i =
2π6
CT T J H12 − 4V1V2 ǫ(Z1, Z2, Z3, X1, X2, X3)
For completeness we can write this same correlator using the notation of Osborn and
where tµνρσα (x) is the OPE structure Tµν (x)Tρσ(0) ∼ x−5tµνρσα J α(0) and X12 = xx12133 −
x223 . The two structures possible, which are linear combinations of the H12 and V1V2, are
1
x512x323x31
3 Iµν µ ′ν′ (x13)Iρσ
ρ′σ′ (x23)tµ ′ν′ρ′σ′α(X12)
(E.2)
(E.3)
(E.4)
(E.5)
1
tµνρσα (x) =
2
tµνρσα (x) =
xγ xµ xρ
4x3 ǫνσαγ + . . . ,
4x
xγ δµρ ǫνσαγ + . . . ,
where the dots represent terms needed to add in ordered for the expression to be symmetric,
traceless and permutation symmetric between the first two pair of indices. The most general
case has t = aˆt1 + ˆbt2. Imposing conservation and comparing with the spinning correlator
formalism we get aˆ = −6ˆb = 3CT T J /π6. Using this information it is straightforward to
apply the same procedure as was done for hT T Oi to obtain the energy matrix elements in
the conformal collider experiment.
F
Computing the bound in the gravity theory
In this appendix we relate the OPE coefficient CT T O to a coefficient, α, in the AdSD
Z √g(R − 2Λ) + (∇χ)2 − m2χ2 + 2αχW 2 ,
Λ = −
(D − 1)(D − 2)
2RA2dS
effective action
S =
MpDl −2
2
where D is the dimension of AdSD. χ is defined to be dimensionless and α has dimensions
In principle we can compute the relation between α and CT T O by computing the
three point function between a scalar and the graviton produced by this cubic term in
the Lagrangian, using Witten diagrams. Instead, we will follow a different route. We will
directly compute the energy correlator in gravity and derive a bound on α by demanding
its positivity. We then relate α and CT T O by demanding that this gravity bound, in terms
of α, matches the bound we obtained in terms of CT T O in the field theory analysis.
We will rely on [
1, 3
] where the energy correlators were computed in gravity. An
important point is that the insertion of T−− corresponds to a shock wave localized in a null
plane. Furthermore, an operator insertion at the origin with definite energymomentum
gives rise to an excitation that crosses this null plane at a localized point. For this reason
the computation of the bound boils down to analyzing the propagation of an excitation
through a suitable gravitational shock wave in flat space. The AdSD space is only relevant
for determining the transverse profile of the shock wave, as we will see below.
For these reasons we consider a shock wave of the form
ds2 = ds2flat + (dx+)2δ(x+)h(y) ,
ds2flat = −dx+dx− + dy2 .
(F.2)
Adding gravitons we get
ds2 = ds2flat + (dx+)2δ(x+)h(y) + dxµ dxνζµ ζνeip.xG(p) + dxµ dxνζ¯µ ζ¯νe−ip.xG¯(p) , (F.3)
with ζ2 = 0, ζµ pµ = 0. Note that the graviton polarization is ζµν = ζµ ζν, and is normalized
to one ζ.ζ¯ = 1. We can think of G(p) and G¯(p) as complex numbers, which in the quantum
theory will be related to a and a†. Inserting (F.3) into (F.1) we can derive the quadratic
HJEP1(207)3
and cubic interaction terms.
2
S =
MpDl −2 Z dx+dx−dD−2y
p+p− + δ(x+)p2−h
G(p)G¯(p) + 4p−p+χ(p)χ¯(p) +
+8p2−αζij∂i∂jhδ(x+)G(p)χ¯(p) + c.c. ,
where we only wrote the terms relevant for our computation, ignoring transverse derivatives
in the kinetic terms. Momentarily setting the scalar field to zero, we see that we have the
following equation for the graviton as it crosses the shock wave
Δhµν ≡ hµν x+=0+ − hµν x+=0− = ip−hhµν .
Exponentiating this, hµν (x+ = 0+) = eip−hhµν (x+ = 0−), we see that the time delay is
simply given by h. This is as expected from (F.2) since we can shift x− by h and make
the term involving h disappear if we ignore its y dependence. So far, we considered the
computation in flat space. An insertion of the null energy integrated along a ray in the
boundary theory gives rise to a shockwave in AdSD which is localized on a null direction.
Its dependence on the transverse directions is the following. The transverse space is an
HD−2. This is easy to see in embedding coordinates where AdSD is W˜ +W˜ − + W µ Wν = −1
(setting RAdSD = 1). The null plane is W˜ + = 0. It contains the null direction parametrized
by W˜ − as well as the transverse space W µ Wµ = −1. The profile of the wave is proportional
to h ∝ (W 0
− W ini)2−D [
1, 3
], with a positive coefficient. Here ~ni is a vector on the sphere
at infinity in the boundary Minkowski space. For (F.4) we need the derivatives at W i = 0,
which are given by
h → h ,
∂i∂jh = [(constant)δij + (D − 2)(D − 1)ninj]h RA2dSD
where the constant does not matter because the graviton is traceless. The relevant
component of the graviton is the one with polarization along ni. This has the expression
1
ζij =
r D − 2
D − 3
ninj − D − 2
δij
(F.4)
(F.6)
(F.7)
The expression for the time delay acting on a superposition of a graviton and a scalar is
now a matrix proportional to
1 γ !
γ 1
,
γ ≡ 4(D − 1)p(D − 3)(D − 2) RA2dSD
,
α
where the matrix is acting on a two dimensional space where one direction is the scalar
and the other is the graviton with polarization (F.7). The unitarity bound comes from the
restriction that the eigenvalues are nonnegative, or γ ≤ 1, which is
α
 
RA2dSD
1
≤ 4(D − 1)p(D − 3)(D − 2)
1
,
where d is the dimension of the boundary. Comparing this with the bound obtained in (3.8),
with the nonEinsteingravity structures set to zero, we obtain (6.2). Of course, once we
get the proportionality constant between α and CT T O for the Einstein gravity case, the
same constant holds also if we add the purely gravitational higher derivative terms that
generate the other tensor structures for hT T T i. We could add them to this computation,
but we expect to reproduce the bounds we got in the general field theory analysis.
F.1
Fourdimensional case
In the special case of the four dimensional theory, we actually have two couplings (6.3).
This leads to a new interaction term in (F.4) of the form
(F.8)
(F.9)
αζij ∂i∂j h → αeζij ∂i∂j h + αoζilǫlj ∂i∂j h ,
where now ǫij is the two dimensional epsilon symbol. This means that the scalar can now
mix with the other graviton polarization component besides (F.7). Namely, defining (F.7)
as ζ⊕, it can also mix with ζij
⊗ ≡ ǫilζlj . Now the time delay is a three by three matrix
⊕
1 γ β
(F.10)
(F.11)
where the rows and columns correspond to the scalar and the two graviton polarizations.
Now the bound is (6.5). Comparing this to (4.13), after setting the nonEinsteingravity
structures to zero, we get the precise mapping to the CT T O coefficients (6.4).
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