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Non-Renormalization Theorems Without Supersymmetry

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Non-Renormalization Theorems Without Supersymmetry CALT-2015-024 Non-renormalization Theorems without Supersymmetry Clifford Cheung and Chia-Hsien Shen Walter Burke Institute for Theoretical Physics California Institute of Technology, Pasadena, CA 91125∗ (Dated: May 11, 2015) We derive a new class of one-loop non-renormalization theorems that strongly constrain the running of higher dimension operators in a general four-dimensional quantum field theory. Our logic follows from unitarity: cuts of one-loop amplitudes are products of tree amplitudes, so if the latter vanish then so too will the associated divergences. Finiteness is then ensured by simple selection rules that zero out tree amplitudes for certain helicity configurations. For each operator we define holomorphic and anti-holomorphic weights, (w, w) = (n − h, n + h), where n and h are the number and sum over helicities of the particles created by that operator. We argue that an operator Oi can only be renormalized by an operator Oj if wi ≥ wj and wi ≥ wj , absent non-holomorphic Yukawa couplings. These results explain and generalize the surprising cancellations discovered in the renormalization of dimension six operators in the standard model. Since our claims rely on unitarity and helicity rather than an explicit symmetry, they apply quite generally. INTRODUCTION where n(A) and h(A) are the number and sum over helic- ities of the external states. Since A is physical, its weight Technical naturalness dictates that all operators not is field reparameterization and gauge independent. The forbidden by symmetry are compulsory—and thus gen- weights of an operator O are then invariantly defined by erated by renormalization. Softened ultraviolet diver- minimizing over all amplitudes involving that operator: gences are in turn a telltale sign of underlying symmetry. w(O) = min{w(A)} and w(O) = min{w(A)}. In prac- This is famously true in supersymmetry, where holomor- tice, operator weights are fixed by the leading non-zero phy enforces powerful non-renormalization theorems. contact amplitude2 built from an insertion of O, In this letter we derive a new class of non- w(O) = n(O) − h(O), w(O) = n(O) + h(O), (4) renormalization theorems for non-supersymmetric the- ories. Our results apply to the one-loop running of where n(O) is the number of particles created by O and the leading irrelevant deformations of a four-dimensional h(O) is their total helicity. For field operators we find: quantum field theory of marginal interactions, ¯ ¯ O Fαβ ψα φ ψα˙ Fα˙ β˙ X ∆L = ciOi, (1) h +1 +1/2 0 −1/2 −1 i (w, w) (0, 2) (1/2, 3/2) (1, 1) (3/2, 1/2) (2, 0) where Oi are higher dimension operators. At leading where all Lorentz covariance is expressed in terms of order in ci, renormalization induces operator mixing via four-dimensional spinor indices, so e.g. the gauge field 2 dci X ¯ (4π) = γ c , (2) strength is F ˙ = Fαβ¯ ˙ + F ˙ αβ. The weights of d log µ ij j ααβ˙ β α˙ β α˙ β j all dimension five and six operators are shown in Fig.1. where by dimensional analysis the anomalous dimension As we will prove, an operator Oi can only be renormal- ized by an operator Oj at one-loop if the corresponding matrix γij is a function of marginal couplings alone. The logic of our approach is simple and makes no ref- weights (wi, wi) and (wj, wj) satisfy the inequalities arXiv:1505.01844v1 [hep-ph] 7 May 2015 erence to symmetry. Renormalization is induced by log wi ≥ wj and wi ≥ wj, (5) divergent amplitudes, which by unitarity have kinematic and all Yukawa couplings are of a “holomorphic” form cuts equal to products of on-shell tree amplitudes [1]. If consistent with a superpotential. This implies a new any of these tree amplitudes vanish, then so too will the class of non-renormalization theorems, divergence. Crucially, many tree amplitudes are zero due to helicity selection rules, which e.g. forbid the all minus γij = 0 if wi < wj or wi < wj, (6) helicity gluon amplitude in Yang-Mills theory. which dictate mostly zero entries in the anomalous di- For our analysis, we define the holomorphic and anti- mension matrix. The resulting non-renormalization the- 1 holomorphic weight of an on-shell amplitude A by orems for all dimension five and six operators are shown w(A) = n(A) − h(A), w(A) = n(A) + h(A), (3) in Tab.I and Tab.II. 1 Holomorphic weight is a generalization of k-charge in super 2 By definition, all covariant derivatives D are treated as partial Yang-Mills theory, where the NkMHV amplitude has w = k + 4. derivatives ∂ when computing the leading contact amplitude. 2 dimension 6 In a renormalizable theory, [g] = 0 or 1, so we obtain dimension 5 F 2φ2 w3, w3 ≥ 2, (9) 6 F 3 F 2φ 2φ3 φ6 4 as a lower bound for the three-point amplitude. 2 F φ 2 2 5 5 2 φ φ Next, consider the four-point tree amplitude. As we F ¯ 2D 2 2 2 3 4 ¯ ¯ φ will see, w4, w4 ≥ 4 for the vast majority of amplitudes. 4 2 φ D The reason is w4 < 4 or w4 < 4 requires non-zero total w 3 ¯2φ2 w F¯2φ2 helicity, which is usually forbidden by helicity selection 2 F¯ ¯2φ rules. To show this, we run through all possible candi- ¯4 ¯2 1 F φ date amplitudes with w4 < 4. Analogous arguments of F¯ ¯2 course apply for w4 < 4. 0 F¯3 1 3 5 Most four-point tree amplitudes with w4 = 1 or 3 w vanish because they have no Feynman diagrams, so 0 2 4 6 w 0 = A(F +F +F ±φ) = A(F +F +ψ±ψ±) Figure 1. Weight lattice for dimension five and six operators, = A(F +F −ψ+ψ+) = A(F +ψ+ψ−φ) suppressing flavor and Lorentz structures, e.g. on which fields + + + − derivatives act. The non-renormalization theorems in Eq. (5) = A(ψ ψ ψ ψ ). let operators to mix into operators of equal or greater weight. Pictorially, this forbids transitions down or to the left. Furthermore, most amplitudes with w4 = 0 or 2 vanish due to helicity selection rules, so 0 = A(F +F +F +F ±) = A(F +F +ψ+ψ−) Since our analysis hinges on unitarity and helicity, the + + + + + resulting non-renormalization theorems are general and = A(F F φ φ) = A(F ψ ψ φ). not derived from an explicit symmetry of the off-shell La- While these amplitudes have Feynman diagrams, they grangian. Moreover, our findings explain the ubiquitous vanish on-shell for their chosen helicities. This leaves a and surprising cancellations [2] observed in the one-loop handful of amplitudes that can in principle be non-zero, renormalization of dimension six operators in the stan- + + + + + + + dard model [3–6]. Lacking an explanation from power 0 6= A(ψ ψ ψ ψ ),A(F φ φ φ),A(ψ ψ φ φ), counting or spurions, the authors of [2] conjectured a hid- for which w4 = 2, 3, 3, respectively. These “exceptional den “holomorphy” enforcing non-renormalization theo- amplitudes” are the only four-point tree amplitudes with rems among holomorphic and anti-holomorphic opera- w4 < 4 that are not identically zero. tors. We show here that this classification simply corre- Since the exceptional amplitudes require external or sponds to w < 4 and w < 4, so the observed cancellations internal scalars, they never arise in theories of only gauge follow directly from Eq. (6), as shown in Tab.II. bosons and fermions, e.g. QCD. The second and third amplitudes require super-renormalizable cubic scalar in- teractions, which we do not consider here. Meanwhile, WEIGHING TREE AMPLITUDES the first amplitude arises from Yukawa couplings of non- holomorphic form, which is to say a combination of cou- We now compute the holomorphic and anti- plings of the form φψ2 together with φψ¯ 2. In a super- holomorphic weights (wn, wn) for a general n-point on- symmetric theory, such couplings would violate holo- shell tree amplitude in a renormalizable theory of mass- morphy of the superpotential. In the standard model, less particles. We start with lower-point amplitudes and Higgs doublet exchange generates an exceptional ampli- then apply induction to extend to higher-point. tude proportional to the product up-type and down-type Working in spinor helicity variables, we consider the Yukawa couplings. This diagram will be important later three-point amplitude with coupling constant g, when we discuss renormalization in the standard model. r3 r1 r2 P h12i h23i h31i , hi ≤ 0 In summary, we find A(1h1 2h2 3h3 ) = g i (7) r3 r1 r2 P [12] [23] [31] , i hi ≥ 0 w4, w4 ≥ 4, (10) which corresponds to MHV and MHV kinematics, |1] ∝ as a lower bound for the four-point amplitude, modulo |2] ∝ |3] and |1i ∝ |2i ∝ |3i. Lorentz invariance fixes the exceptional amplitudes. P P Finally, consider a general higher-point tree ampli- the exponents to be ri = −ri = 2hi − h and i ri = P tude, A , which on a factorization channel degenerates i ri = 1 − [g] by dimensional analysis [7]. According i to Eq. (7), the weights of the three-point amplitude are into a product of amplitudes, Aj and Ak, where P i X h −h (4 − [g], 2 + [g]), i hi ≤ 0 fact[Ai] = Aj(` )Ak(−` ), (11) (w3, w3) = P (8) `2 (2 + [g], 4 − [g]), i hi ≥ 0 h 3 Oi is radiatively generated by Oj. In practice, γij is factorize loop extracted from the one-loop amplitude Ai involving an insertion of Oj that has precisely the same external wi = wj + wk 2 − states as the tree amplitude Ai involving an insertion Ai Aj Ak loop of Oi.
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