Horizon exchange and the universal all-spin response of Kerr

Introduction

The mass of a Kerr black hole can be written in terms of its irreducible mass $m_{\mathrm{ir}}$ and angular momentum $J$ as [1, 2]

It is invariant under

which exchanges the outer and inner horizon areas. This algebraic exchange is not a symmetry of black-hole scattering. The physical solution is selected by an ingoing boundary condition at the event horizon, whereas the inner horizon lies behind it. Nevertheless, as we show, the radial perturbation equation is invariant under the exchange. It therefore maps the physical solution to a well-defined image solution, and the two associated tidal responses obey an exact relation.

This comparison has a direct physical use. Black-hole perturbation theory (BHPT) has produced small-spin Kerr-Compton data below extremality and continuation-dependent coefficients above extremality [3, 4, 5, 6, 7]. BHPT data through sixth order in spin are already used in the aligned-spin analysis of Ref. [6], and recently reassembled in Ref. [7]. We give an exact relation that explains how the two horizon-selected responses are related, together with its general-multipole consequence in the long-wavelength regime with the black-hole spin left unconstrained. The far-zone propagation is common to the two solutions and all dependence on which horizon condition is imposed is carried by the near-zone response. In the low-frequency quadrupole, the exchange-odd term changes the scattering probability and gives a definite interference contribution to physical absorption. The exchange-even response contains the conservative phase and contact information, but need not be purely conservative: it can also contain exchange-even absorption, in particular in the superradiant regime.

Horizon exchange and response

For Kerr, write $r_\pm= M(1 \pm\kappa)$, with $a = J/M$ the specific angular momentum, $q = a/M$ the dimensionless spin and $\kappa= \sqrt{1-q^2}$. Exchanging the horizons is $\kappa\mapsto-\kappa$. The radial Teukolsky equation [8],

where $m$ is the azimuthal mode number and $K=(r^2+a^2)\omega-am$, depends on the horizons only through $M$ and $a$ and is consequently invariant. The local horizon exponents are governed by

with $\Omega_\pm=a/(2Mr_\pm)$ and $T_\pm=(r_\pm-r_\mp)/(8\pi Mr_\pm)$. The angular velocity of the event horizon is $\Omega_H:=\Omega_+$ and $T_H:=T_+$ is its Hawking temperature. Thus $\tau$ changes sign under the exchange. The physical solution $R_+$ is ingoing at $r_+$ and its image $R_-^{(\eta)}$ is selected by the corresponding ingoing local exponent at $r_-$. At large radius write each solution in the MST Floquet basis $R_C^\nu,R_C^{-\nu-1}$ [9, 10, 11]:

The ratio $\mathcal{K}_+$ is the MST near-zone (tidal-response) coefficient selected by the physical event-horizon boundary condition, while $\mathcal{K}_-^{(\eta)}$ is the corresponding coefficient for the auxiliary inner-horizon image. Here and below, “near zone” and “far zone” refer operationally to the two factors in the MST factorization of the partial-wave $S$ matrix, not to a sharp division of the radial coordinate. The near-zone factor is the response-dependent factor containing $\mathcal{K}$, which carries the information about the horizon boundary condition. The far-zone factor is the remaining propagation and asymptotic normalization factor, common to the two horizon-selected solutions. In a low-frequency matched-asymptotic interpretation, these factors encode, respectively, the horizon-sensitive response and propagation from the matching region to infinity. This terminology follows the factorization used in Ref. [12].

Exact exchange law. —Our main result is the following. For integer spin weight $s$, generic frequency ($2\nu\notin\mathbb{Z}$ and $\epsilon_\pm\ne0$), and an image solution continued from $r_-$ to infinity along a path labelled by $\eta=\pm1$, the Floquet exponent and far-zone scattering data are exchange-invariant, while the near-zone responses satisfy

The two values of $\eta$ specify the two homotopy classes of continuation around the outer horizon (Figure 1). They are path labels, not black-hole parameters. They correspond to the two analytic continuations of the inner-horizon solution and should not be confused with the $\alpha$ bookkeeping labels used for particular contact-term contributions in the Kerr-Compton literature. The ratio is independent of the normalization of either radial solution or of the common Floquet basis. Equation (6) is exact in frequency and follows from the $2\times2$ monodromy algebra, not from a low-frequency expansion or an explicit connection formula.

To see the origin of the relation, let $M_\pm$ denote the monodromies around the two horizons and $M_\infty$ the monodromy enclosing both, based at a point $r_0>r_+$ [Figure 1(a)]. In the Floquet basis, $M_\infty=\operatorname{diag}(\lambda,\lambda^{-1})$, with $\lambda=e^{2\pi i\nu}$. The physical solution and its outgoing partner are eigenvectors of $M_+$ with eigenvalues $a^{\pm1}$, where $a=e^{2\pi\epsilon_+}$. The image is an eigenvector of $M_-$ with eigenvalue $b=e^{2\pi\epsilon_-}$. Using $M_\infty=M_-M_+$ or $M_+M_-$, the trace condition $\operatorname{tr}M_-=b+b^{-1}$ fixes the second eigenvector of $M_+$. The resulting image eigenvector gives

which is Eq. (6), since $a/b=e^{2\pi\tau}$. The full proof, including conventions and path ordering, is provided in Supplemental Sec. S1 [13].

At zero frequency, $\nu=\ell$ and the exchange law becomes trivial. This is consistent with the polynomial truncation of the static perturbation problem and with the vanishing of the four-dimensional static Kerr Love numbers established in Refs. [14, 15, 16, 17, 18]. The triviality of the monodromy relation in this limit admits a consistency observation. It is not, by itself, a matching calculation or a proof that the corresponding effective worldline Wilson coefficients vanish. In the EFT description, static Love numbers are defined by matching the low-frequency scattering amplitude after separating universal long-range contributions. Recent amplitude-based matching calculations confirm the vanishing of the static Schwarzschild Love numbers on shell [19, 20]; these Schwarzschild results provide a complementary EFT perspective, rather than an independent derivation of the rotating Kerr result. The non-trivial dynamics and the running of the dynamical tidal response are controlled by $\delta=\nu-\ell=O(\epsilon^2)$ [21, 22]. The MST expression for the partial-wave scattering matrix can be written

where $F_{\ell m}$ is the horizon-independent far-zone propagation and asymptotic-normalization factor. The rational factor is the near-zone response factor and depends either on $\mathcal{K}_+$ or on $\mathcal{K}^{(\eta)}_-$. The label “far zone” here refers to this factor in the MST partial-wave factorization, not to the complete point-particle Compton amplitude or the complete covariant contact polynomial.

Universal fixed-$\omega$ response at arbitrary multipole. —The exact law (6) has a particularly simple fixed-$\omega$ limit. Take $\epsilon=2M\omega\to0$ at fixed $c=a\omega$, and write $\delta=\nu-\ell$ and $u=i\tau+m$ where $\tau$ is defined in (4), both of order @page 2/14 starts_mid=1 ends_mid=1 $\epsilon^2$ Define the near-zone factor by $\mathcal{N}^{\mathrm{NZ}}_{\ell m}(\eta)=1+H^{\mathrm{even}}_{\ell m}+\eta H^{\mathrm{odd}}_{\ell m}$, so that $H^{\mathrm{odd}}_{\ell m}$ is one half of the difference between the two continuation choices. The exact exchange law then fixes its leading term for every integer $\ell$ and $|m|\leq\ell$:

where

and $R_{\ell m}(c):=\lim_{\epsilon\to0}\mathcal{X}_{-\nu-1}/\mathcal{X}_{\nu}$, where the limit is taken at fixed $c=a\omega$ (Supplemental § 1.4), with $R_{\ell m}(0)=1$. Equation (9) is the all-orders-in-$a\omega$ result at leading order in $M\omega$, not a finite-frequency formula for the complete scattering amplitude. Its important content is that the dependence on the anomalous dimension $\delta$ cancels between the monodromy relation and the MST near-zone factor. The remaining function $R_{\ell m}$ is determined by the limiting MST recurrence (Supplemental § 1.4). For gravitational perturbations ($s=-2$), its quadrupolar expansion begins

Equation (9) is the main all-multipole result of the paper in the fixed-$a\omega$, leading-small-$M\omega$ regime: it gives the leading exchange-odd response to all orders in $a\omega$, with the entire multipole dependence encoded in the limiting MST series ratio $R_{\ell m}$. The quadrupolar Bose-factor expression below and the polar closed-form amplitude (23) are special projections of it.

Thus the far-zone contribution cancels when the two horizon-selected amplitudes are subtracted, and the difference is sourced by the change in the near-zone response. Define $S^-_{\eta}=[S(\mathcal{K}_+)-S(\mathcal{K}^{(\eta)}_-)]/2$ and expanding the near-zone factor to first order in $\mathcal{K}_+$ gives

Here $E$ is the linear coefficient of the near-zone factor in the MST $S$ matrix. The far-zone factor $F$ is needed to normalize the full scattering amplitude, but it carries no horizon-exchange odd information as this is provided by the factor $E(\mathcal{K}_+-\mathcal{K}^{(\eta)}_-)$ which is the leading change in the scattering amplitude caused by replacing the event-horizon condition by its inner-horizon image. Near $\nu=\ell$ both $E$ and $1-\mathcal{K}^{(\eta)}_-/\mathcal{K}_+$ vanish linearly in $\delta=\nu-\ell$, while $\mathcal{K}_+$ has a double pole. Therefore, the $E(\mathcal{K}_+-\mathcal{K}^{(\eta)}_-)$ starts at $O(\epsilon^{2\ell+2})$. This is the lifting of the static polynomial-truncation degeneracy. This linearization is used only for the low-frequency quadrupolar extraction below. The all-orders-in-$a\omega$ result (9) in the leading-small-$M\omega$ regime is instead obtained from the full MST near-zone factor and the exact exchange law.

For gravitational quadrupole perturbations ($s=-2$, $\ell=2$), the leading term can be evaluated analytically. With $\varpi=m\Omega_H/(2\pi T_H)=mq/\kappa$, for the continuation $\eta=+1$, and after removing the common long-range phase, we obtain

which is real at this order. The horizon Bose factor makes the thermal weighting of the horizon response explicit. In the unit-flux partial-wave convention (see the Supplemental Sec. S8), we use the parity-resolved reflection coefficient $S_P$ for flux balance. The helicity-preserving coefficient $S$ used in the Compton amplitude differs by the helicity-reversing channel. Their conversion is exchange-even, so it leaves the leading exchange-odd term below unchanged. The interference of the exchange-odd term with the exchange-even amplitude gives

where $S^{\mathrm{even}}=1+O(\epsilon)$ after the common phase is removed. Here $\Delta_{\mathrm{odd}}\mathcal{P}_H$ denotes the term linear in the exchange-odd amplitude in the expansion of $1-|S_+|^2$. This is an interference contribution, not a separately measurable absorption probability. The full absorption probability also contains the exchange-even contribution and, beyond the helicity-preserving normalization, the helicity-reversing channel described in Supplemental Sec. S8. For $\eta=-1$ the Bose factor $\bar{n}=1/(e^{2\pi\varpi}-1)$ is replaced by $\bar{n}+1$, and the difference, $(\pi/225)m q[1+(m^2-1)q^2][4+(m^2-4)q^2]\epsilon^6$, is polynomial in the spin. In the Schwarzschild limit this exchange-odd contribution is $2\epsilon^6/225$, exactly one half of Page’s quadrupolar absorption probability $4\epsilon^6/225$ [23], with the other half provided by the exchange-even sector. For $m=0$ the exchange-odd share is $\kappa/(1+\kappa)$. This is a useful quantitative illustration that horizon-exchange parity is not identical to the conservative/absorptive split. More generally, the exchange-even sector can also change the modulus of $S_{\ell m}$. In particular, the leading superradiant term $2q(1+3q^2)\epsilon^5/225$ of $|S_{22}|$ is exchange-even [21]. The exchange parity is therefore a classification of the response under the auxiliary horizon map, not a synonym for the real/imaginary or conservative/absorptive decomposition at all orders.

The physical interpretation is constrained by the standard flux balance. For the event-horizon solution, the Teukolsky–Starobinsky identities convert a conserved cross-spin radial Wronskian into $\mathcal{F}^{\infty}_{\mathrm{in}}-\mathcal{F}^{\infty}_{\mathrm{out}}=\mathcal{F}_H$, with $\mathcal{F}_H$ proportional to $(\omega-m\Omega_H)$ times a non-negative horizon-amplitude norm in a flux-normalized convention [24]. Thus the horizon energy flux changes sign at the superradiant threshold. This does not make $S^{-}$ a difference of two horizon fluxes because the inner-horizon image is not a second solution in the physical exterior scattering problem, and the flux is quadratic in the perturbation. Rather, $S^{-}$ measures the change in the scattering amplitude induced by changing the local horizon condition. Its real leading term changes the physical scattering probability and hence contributes to the flux balance. The next term in the low-frequency expansion and the unit-flux normalization are given in Supplemental Secs. S6 and S8.

Conservative response and zeta values. —The exchange-even projection contains a contribution from the far zone as well as the near-zone response. Combining them, the exchange-even quadrupolar coefficient at order $\epsilon^6$ is

where $P_{5}(q)$ denotes a degree-five polynomial and

The digamma and logarithmic terms are derived analytically from the exchange law and the MST representation. The polynomial part is identified numerically [25] with its continuation tested against independent MST data on the superextremal branch (Supplemental Sec. S4). The logarithmic coefficient describes the running of the dynamical Love response, while the digamma term gives its non-polynomial spin dependence. In this sense the exchange construction does not identify conservative physics with a formal even projection. It supplies the horizon-sensitive piece, which must be combined with the invariant far-zone phase to obtain the conservative response.

The separation into real and imaginary parts of the digamma function explains the transcendentality pattern:

Thus the conservative spin coefficients contain odd zeta values, whereas the absorptive coefficients contain even zeta values. This transcendental pattern and the relevant low-spin coefficients have antecedents in the BHPT and Kerr-Compton literature [6, 7]. The new statement is that the horizon-exchange law organizes the two structures as projections of the same near-zone response.

Connection across extremality. —For incidence along the spin, only $m=2$ contributes. The digamma contribution to the quadrupolar conservative contact coefficient admits the all-orders-in-spin expression

recalling that $\kappa=\sqrt{1-q^{2}}$. Its low-spin expansion reproduces the BHPT coefficients of [6], and the all-orders resummation in (19) is the new result. Analytically continuing the symmetrized digamma representation across $q=1$ gives, for $q>1$,

The pole of the second digamma function at argument $-1$ controls the large-spin limit,

This reproduces the corresponding $\alpha$-tagged coefficient of Ref. [4]. Here $\alpha$ is the bookkeeping label in that amplitude decomposition for terms associated with the continued digamma contribution. Superextremality itself is specified by $q>1$. The poles of the continued digamma reproduce the corresponding $\alpha$-tagged terms at fifth and sixth order in spin, while Eq. (13) reproduces the continuation-odd coefficients. The branch prescription, pole structure, and large-$q$ derivation are given in Supplemental Sec. S4.

Seventh order and beyond. —At fixed $a\omega$, the exchange law also determines the leading small-$M\omega$ exchange-odd response for general $\ell$. Combining this result for $\ell=2$ with the spheroidal-harmonic mixing and the leading $\ell=3$ contribution gives, for incidence along the spin, the first three terms of

where $\Theta$ is the scattering angle and $f_{0}$ the spinless amplitude. The first two terms reproduce the sixth-order result, Eq. (3.52) of Ref. [4] for the superextremal analytic continuation, and the $a^{7}$ and $a^{8}$ coefficients (second and third lines) agree with Eq. (5.5) of Ref. [12], which extends this to eighth order in spin. For this configuration the higher orders are fixed as well. The limiting amplitude computed independently from the full MST partial-wave sum ($\ell\le19$, 110 digits) is reproduced order by order through $O(a^{13})$, to relative accuracy $10^{-15}$, by the closed form

with $c=a\omega$ and $v=\sin^2(\Theta/2)$. Its expansion confirms the $a^7$ coefficient and gives the $a^8$ term on the third line of Eq. (22) with the next orders listed in Supplemental Sec. S3. Equation (23) was identified from high-precision MST data and independently checked against the segment representation. Its all-order closed form is therefore a strongly validated reconstruction, rather than a consequence already proved by the monodromy algebra alone. In the formal $M \to0$ limit at fixed $a$ and $\omega$, the horizon locations approach $r=\pm ia$, and the exchange-odd parts of the partial-wave $S$-matrix elements are given exactly by an integral along the segment joining them plus an endpoint term [Supplemental Eq. (40)], which we checked against MST to $10^{-16}$. Correspondingly, Eq. (23) is a superposition of $e^{-2a'\omega v}$, the exponential of Ref. [4] for spin $a'$ at this kinematics, with $a'$ ranging over $[-a,a]$ [Supplemental Eq. (41)].

Discussion. —The exact response law (6) is the structural theorem, and Eq. (9) is its principal physical consequence giving a universal all-orders-in-spin expression for the leading exchange-odd response at every multipole. Horizon exchange is not a symmetry between two physical Kerr scattering experiments because the inner-horizon solution is an auxiliary solution of the same radial equation, selected by the exchanged local boundary condition. The far-zone propagation is common to the two solutions, so their difference isolates the change in the near-zone response. At leading quadrupolar order this difference is real after removal of the common long-range phase and therefore contributes to the physical scattering probability and flux balance. It is an interference contribution, not a separate horizon-flux probability. The exchange-even sector contains the conservative phase/contact information but can also contain absorption, including the leading superradiant term. Horizon-exchange parity is therefore not identical to the conservative/absorptive decomposition.

The digamma response organizes the previously computed subextremal and superextremal Kerr-Compton coefficients. The superextremal regime is the analytic continuation $q>1$ with imaginary $\kappa$, this is not represented by setting a contact-term bookkeeping parameter to a particular value. In particular, the far-zone point-particle amplitude and its covariant contact polynomial must be distinguished from the near-zone response constrained by Eq. (6). The polar closed form above fixes a polar projection only, not every independent far-zone contact coefficient. We reproduce the digamma origin of the non-rational spin dependence [7], the $\alpha$ label on a digamma function [12], and the monodromy approach to Kerr scattering coefficients [26, 27]. The new results derived here are the exact law and its general-multipole consequence (9) in the fixed-$\alpha$, leading-small-$M\omega$ regime, together with the exchange-parity organization it induces. The exact law, the general-multipole formula, and the leading quadrupolar exchange-odd response are analytic results. The polynomial part of the conservative coefficient is a numerical identification supported by high-precision continuation. Extending the closed-form analysis to higher multipoles, deriving the remaining polynomial terms analytically, and deriving Eq. (23) by summing the segment representation over multipoles are important next steps.

Acknowledgments. —We thank F. Bautista for feedback and discussions that led to improvements in the manuscript. This work has been supported by the Agence Nationale de la Recherche (ANR) under grant Observables (ANR-24-CE31-7996). We acknowledge the use of Claude, an AI model developed by Anthropic, to proofread the text and to write and run the Python code for the numerical checks quoted in the Supplemental Material. The author takes full responsibility for the content of the paper.

Code availability.—The scripts and data reproducing the numerical checks of this paper are available at [25].

Supplemental Material

The Supplement establishes the monodromy exchange law, derives its fixed-$a\omega$, small-$M\omega$ and low-frequency consequences, and documents the numerical reconstruction and flux normalization. Each equation is introduced where its definitions and derivation are discussed. All scripts and data are available at [25].

S1. The exchange law: setting and proof.

Setting. For integer $s$ the radial Teukolsky equation has regular singular points at $r=r_{\pm}$, with exponents $\{-s-i\epsilon_{\pm},i\epsilon_{\pm}\}$, and an irregular singular point at $r=\infty$, as in the monodromy approach to Kerr scattering [26, 27]. Fix a base point $r_{0}>r_{+}$ on the real axis. The monodromy $M_{\infty}$ along a large circle through $r_{0}$ has eigenvalues $\lambda^{\pm1}$, $\lambda=e^{2\pi i\nu}$, with eigenfunctions the Floquet solutions $R_{C}^{\nu}$, $R_{C}^{-\nu-1}$; for $2\nu\notin\mathbb{Z}$ these are unique up to their normalizations, and in this basis $M_{\infty}=\operatorname{diag}(\lambda,\lambda^{-1})$. Let $M_{+}$ be the monodromy along a loop from $r_{0}$ around $r_{+}$ alone. The physical solution $R_{+}$, defined near $r_{+}$ by its exponent $-s-i\epsilon_{+}$, is the eigenvector of $M_{+}$ with eigenvalue $a=e^{2\pi\epsilon_{+}}$, and the outgoing solution at $r_{+}$ is the eigenvector with $a^{-1}$; for integer $s$, $\det M_{+}=1$, and $a\ne a^{-1}$ for $\epsilon_{+}\ne0$. The image $R_{-}$ is defined near $r_{-}$ by the exponent $-s-i\epsilon_{-}$ and continued to $r_{0}$ along a path passing $r_{+}$ on one side. The corresponding loops around $r_{-}$, with tails passing $r_{+}$ above ($\eta=+1$) or below ($\eta=-1$), give two monodromies $M_{-}^{(\eta)}$ (Fig. 1). With the monodromy of a path $\gamma$ followed by $\gamma'$ equal to $M_{\gamma'}M_{\gamma}$, so that monodromies act on the coefficient vectors of solutions in the Floquet basis, $M_{\infty}=M_{-}^{(-)}M_{+}=M_{+}M_{-}^{(+)}$; hence $M_{-}^{(+)}=M_{+}^{-1}M_{-}^{(-)}M_{+}$, and the two images differ by the action of $M_{+}$. $R_{-}^{(\eta)}$ is the eigenvector of $M_{-}^{(\eta)}$ with eigenvalue $b=e^{2\pi\epsilon_{-}}$.

Convention independence. Writing each solution as $R\propto R_{C}^{\nu}+\mathcal{K}R_{C}^{-\nu-1}$, a rescaling $R_{C}^{\nu}\to c_{1}R_{C}^{\nu}$, $R_{C}^{-\nu-1}\to c_{2}R_{C}^{-\nu-1}$ maps $\mathcal{K}\to(c_{1}/c_{2})\mathcal{K}$ for every solution, so $\mathcal{K}_{-}^{(\eta)}/\mathcal{K}_{+}$ is invariant. The normalizations of $R_{\pm}$ do not enter $\mathcal{K}$. The relabelling $\nu\to-\nu-1$ exchanges the two Floquet solutions, maps $\mathcal{K}\to1/\mathcal{K}$ and $\lambda\to\lambda^{-1}$, and maps the right-hand side of Eq. (6) to its inverse. Equation (6) is therefore a statement about the pair $(R_{+},R_{-}^{(\eta)})$ alone.

Proof. With $M_{+}=P\operatorname{diag}(a,a^{-1})P^{-1}$, $P=\begin{pmatrix}1&1\\ \mathcal{K}_{+}&Y\end{pmatrix}$, the condition that $M_{-}^{(-)}=M_{\infty}M_{+}^{-1}$ have eigenvalues $b^{\pm1}$ (equivalently $\operatorname{tr}M_{-}^{(-)}=b+b^{-1}$) gives

and the $b$-eigenvector $(1,\mathcal{K}_{-}^{(\eta)})$ has $\mathcal{K}_{-}^{(\eta)}/\mathcal{K}_{+}=-\lambda^{-1}(\lambda a-b)/(\lambda b-a)$. For $M_{-}^{(+)}=M_{+}^{-1}M_{\infty}$ one finds $-\lambda(\lambda a-b)/(\lambda b-a)$; the two differ by $\lambda^{2}$. With $a=e^{2\pi\epsilon_{+}}$, $b=e^{2\pi\epsilon_{-}}$, $a/b=e^{2\pi\tau}$ and $\lambda=e^{2\pi i\nu}$, $(\lambda a-b)/(\lambda b-a)=\sin\pi(\nu-i\tau)/\sin\pi(\nu+i\tau)$, which gives Eq. (6). Half-integer $s$, for which $\det M_{\pm}=-1$, is not covered.

S2. Fixed-$a$, small-$M$ limit and limiting MST recurrence.

With the $n=0$ terms of the MST series made explicit [12],

where $\mathcal{X}_{\nu}$ are ratios of MST series, invariant under the exchange. The function entering Eq. (9) is $R_{\ell m}(c)=\lim_{\epsilon\to0}\mathcal{X}_{\nu-1}/\mathcal{X}_{\nu}$ at fixed $c=a\omega$, normalized by $R_{\ell m}(0)=1$. For $\epsilon\to0$ at fixed $a\omega$, $2\epsilon\kappa\to4ia\omega$, $\nu=\ell+\delta$ and $i\tau=-m+u$ with $\delta,u=O(\epsilon^{2})$. Using $1/\Gamma(-N-z)\simeq(-1)^{N+1}Nz$,

and $e^{i\pi\nu}+e^{-i\pi\nu}\rho\simeq(-1)^{\ell}2i\delta/\epsilon$. Thus the leading small-$M\omega$ exchange-odd response, for general integer $\ell$, is

where $\mathcal{N}_{\ell m}=(\ell+m)!(\ell-m)![(\ell-s)!(\ell+s)!]^2/[(2\ell)!(2\ell+1)!]^2$ and $R_{\ell m}$ is the limiting ratio of MST series at fixed $a\omega$. This is the expression used in the leading-small-$M\omega$ amplitude calculation below.

Limiting recurrence. At $\epsilon\to0$, $c=a\omega$ fixed, the MST coefficients become

with $E(c)$ the spin-weighted spheroidal eigenvalue. For $s=-2$, $\ell=2$ the negative-$n$ coefficients drop out and

with $a_1=-\frac{5}{9}(3+m)c+\frac{5m}{81}(3+m)c^2$ and $a_2=\frac{5}{49}(3+m)(4+m)c^2$. Expanding the resulting series ratio gives, for $s=-2$ and $\ell=2$,

which supplies the quadrupolar correction entering Eq. (27) and the amplitude in § 1.5.

$\eta$-odd amplitude and higher orders.

The helicity-preserving amplitude is

with $S_{\ell m}={}_{-2}S_{\ell m}$ at $c=a\omega$ [28]. Through $O(a^6)$ only $\ell=2$ contributes to $f^{\mathrm{odd}}$, and with $f_0=M\cos^4(\Theta/2)/\sin^2(\Theta/2)$ (frame phase restored by the addition theorem)

$Y_{\ell m}={}_{-2}Y_{\ell m}$. At $O(a^7)$ the numerator acquires $-\frac{(a\omega)^2}{3969}\sum_m(-1)^m(3+m)!(3-m)!Y_{3m}^{*}Y_{3m}$. For $s=-2$, $m=2$ and $x=\cos\Theta$,

and

which give Eq. (22) through $O(a^7)$.

Independent check and higher orders. We computed the polar limiting amplitude directly in the fixed-$a\omega$, small-$M\omega$ regime: $\epsilon=10^{-16}$ at fixed $c$, full MST (near and far zone) at 110 digits for $\ell=2,\ldots,19$ and both $\eta$, 28 Chebyshev nodes $|c|\leq0.2$, projected onto the polar harmonics. Through $O(a^{13})$ it agrees, to $10^{-15}$ relative ($3\times10^{-16}$ through $O(a^{10})$), with

For $\eta=+1$ this is the exponential $e^{-2cv}$, and the $\eta$-odd part is Eq. (23). With $Q(b)=b^3-3b^2+6b-6$ the integral is elementary,

Its expansion,

is Eq. (22) with $\cos\Theta=1-2v$. Writing $f^{\mathrm{odd}}/f_0=-\frac{64}{15}vc^5\sum_{k\geq0}b_k(x)c^k$, with $b_0,\ldots,b_3$ read off from the bracket of Eq. (22), the orders tested through $O(a^{13})$ are completed by

The odd-order coefficients contain the factor $3+x$: at $x=-3$ ($v=2$) Eq. (23) is odd in $c$. This is a statement about the polar projection of the exchange-odd amplitude only. It does not determine the complete covariant far-zone contact polynomial: independent contact structures can vanish in the polar limit. This is consistent with the polar simplification in Ref. [29], where the additional contact deformation is not resolved by that kinematic slice.

Segment representation. The expansion at fixed $a$ and $\omega$ starts at first order in $M$. Expanding the radial operator of the main text as $T_M=T_0+MT_1+O(M^2)$ gives, for $s=-2$,

At $M=0$ the radial equation is the spheroidal equation. Its solution regular at both $r=\pm ia$ is $R_0(r)=\Delta_0(r)\widehat{S}_{\ell}(ir/a)$, where $\widehat{S}_{\ell}(x)$ is $S_{\ell2}$ as a function of $x=\cos\theta$, continued to complex $x$ (it is entire). In this limit $r_\pm\to\pm\eta ia$, so the continuation $\eta$ imposes the ingoing condition at $r=\eta ia$. The local exponents are $\{3,-1\}$ at $r=ia$ and $\{1,1\}$ at $r=-ia$, and the ingoing ones are shifted by $-iM\omega$. Writing both solutions as $R_0+MR_1^{(\eta)}$ and integrating the Green identity of the formally self-adjoint operator $\Delta^sT_M$ from one horizon to the other gives the $\eta$-odd part of the partial-wave element,

where $B^{\mathrm{inc}}$ and $B^{\mathrm{ref}}$ are the coefficients of $r^{-1}e^{-i\omega r}$ and $r^3e^{i\omega r}$ in $R_0$. The integrand has a double pole without residue at $r=-ia$, and FP denotes its finite part. The last term is the boundary term at the degenerate horizon $r=-ia$; its $2c$ comes from the ingoing exponent, and the boundary term at $r=ia$ vanishes. On the segment the radial function is the angular one, $R_0(ia')=(a^2-a'^2)\widehat{S}_\ell(-a'/a)$, and the endpoint term involves only $\widehat{S}_\ell$ at $\theta=0$. Equation (40) reproduces the MST values of $e^{2i\delta_{\ell2}}\big|_{\mathrm{odd}}$ computed above, for $\ell=2,\ldots,6$ at two values of $c$, to $4\times10^{-7}$, the size of the finite-$\epsilon$ systematic. It includes the terms nonlinear in $K$ dropped in Eq. (12), which are not small here: since $u/\delta=O(c^{-2})$ in Eq. (26), the exchange-even part of $\mathcal{K}$ is enhanced, and these terms change the $\ell=2$ element already at relative order $c^3$. Equation (23) can be rewritten as

a superposition of the exponential $e^{-2a'\omega}$ for spin $a'=(2t-1)a\in[-a,a]$, which suggests identifying $t$ with the position $ia'$ on the segment. Likewise $f^{(-)}/f_0=e^{2cv}/(1-v)^4+e^{2c(2-v)}[Q(4c(1-v))/(1-v)^4-Q(4c)]/6$ contains the exponential for spin $-a$, matching the endpoint $r=-ia$. Summing Eq. (40) over multipoles would derive Eq. (35); we have not carried this out.

Low-frequency expansion and numerical reconstruction.

Continuation of the all-orders digamma expression. For subextremal real $q$ we have $\kappa=\sqrt{1-q^2}>0$. The real part in Eq. (19) can be written as

The right-hand side defines the meromorphic representative to be continued in complex $q$, away from the digamma poles. Continuing around the branch point at $q=1$ gives $\kappa=\pm i\lambda$, where $\lambda=\sqrt{q^2-1}$ on the superextremal real axis. The two digamma terms are exchanged when the sign of $\kappa$ is reversed, whereas $\log\kappa=\log\lambda\pm i\pi/2$. Taking the real part on this axis therefore gives Eq. (20). The continuation is understood along paths in the complex $q$-plane that avoid the poles. On the real superextremal axis these occur when

or, equivalently,

For completeness, the large-$q$ behavior follows from

Using the digamma expansion $\psi(-1+\delta)=-1/\delta+O(1)$ as $\delta\to0$, we obtain

The other digamma term and the logarithm grow only logarithmically. Consequently the square bracket in Eq. (20) is $q^2/2+O(\ln q)$, which yields Eq. (21). This is the $\alpha$-tagged $q^5$ coefficient identified in Ref. [4]. The label $\alpha$ there is an amplitude-decomposition bookkeeping label, not a Kerr parameter and not a label for either side of extremality. In particular, it should not be confused with the physical condition $q>1$ or with the path label used for the continuation.

The physical solution $S_{2m}$ was computed from the MST continued fraction at 90-digit precision, using 12 angular nodes at each of $r=2\times10^{-5}$ and $10^{-5}$ in $\epsilon$, and 40 ($m=2$) or 24 (otherwise) nodes on $|q|=0.3$. Both $\kappa$ and $-\kappa$ were evaluated, with $\log(-\kappa)=\log\kappa+i\pi$; the spheroidal eigenvalue was computed at complex $c=q\epsilon/2$. The fitted coefficient at order $\epsilon^6$ is

where $p_5^{(m)}(q)$ is the degree-five polynomial identified numerically; its exchange-odd and exchange-even parts are Eqs. (13) and (15).

The logarithmic coefficient is extracted by a joint two-radius fit of the Fourier coefficients at orders $n=6,7,8$, which removes leakage from higher powers times $\ln\epsilon$. A fit to $c_6$ alone determines $B_6$ only to $5\times10^{-5}$–$3\times10^{-4}$, whereas the joint fit gives $B_6=i\Pi_m/450$ to $10^{-10}$ for every $m$. The exchange-odd terms have no logarithm through $\epsilon^7$. Independent free fits recover the coefficients of $\Pi_m\ln\kappa$ and $\Pi_m\operatorname{Re}\psi$ as $-1/450$ to $3\times10^{-9}$. After subtracting Eq. (47), fitted Taylor coefficients for $6\le p\le20$ are below $2\times10^{-11}$, while retained coefficients are $O(10^{-5}$–$1)$.

The polynomial determined at $|q|\le0.3$ also predicts the super-extremal branch. For $m=2$, continuation of Eq. (15) agrees with direct MST to $4\times10^{-12}$ for $1.5\le q\le30$ and to $3\times10^{-10}$ for $50\le q\le5000$ ($6\times10^{-13}$ at $q=5000$). This bounds an additional $q^6$ coefficient by $3\times10^{-18}$, compared with $2\times10^{-11}$ from the direct Taylor fit; higher powers are more strongly constrained. The $q^5$ coefficient of $p_5^{(2)}$, $-2.3457997302\times10^{-5}i$, agrees to $10^{-15}$ with the exactly known $\eta=+1$ limiting value divided by 32. For $q>1$, MST values were taken at fixed $q$ on circles of radius proportional to $1/q$ (three radii for $q\ge50$).

Contact polynomial and general kinematics.

Following Refs. [4, 5], write $f=f_0(e^{(2\omega+k_3-k_2)\cdot a}+P_\xi)$, where

and

(terms vanishing through $O(a^6)$ omitted). In the rest frame, for all-incoming $k_2=\omega(1,\hat n_{\rm in})$, $k_3=-\omega(1,\hat n_{\rm out})$ and mostly-minus metric, $\xi^{-1}=-\sin^2(\Theta/2)$, and

The far-zone factor in the partial-wave factorization is not itself the complete covariant contact polynomial $f^{\rm FZ}$: the latter is obtained by matching the far-zone partial waves to a covariant amplitude at generic kinematics. A polar restriction cannot fix every independent coefficient, since contact structures that vanish in that limit remain unconstrained. The BHPT side is $\sum_m h_m S^*_{2m}S_{2m}$, normalized as in § 1.5; $h_m$ is the pole part of the $\ell=2$ element:

At random kinematic points the ratios to the $\alpha$-tagged ansatz are constant to $10^{-30}$, whereas the opposite helicity fails. A free fit of the full $\eta$-odd amplitude reproduces all nine $\eta$-odd coefficients with residual $<10^{-23}$; see Table S1.

	$j=0$	$j=1$	$j=2$
$c_3^{(j)}$	$a^5$, even	$\frac{64}{15}$	$\frac{16}{15}$
$c_4^{(j)}$	$a^5$, odd	$\frac{64}{15}$	$\frac{16}{5}$
$c_6^{(j)}$	$a^6$, even	$\frac{128}{45}$	$\frac{32}{9}$
$c_8^{(j)}$	$a^6$, even	$-\frac{512}{45}$	$-\frac{160}{9}$
$c_9^{(j)}$	$a^6$, odd	$-\frac{128}{45}$	$-\frac{32}{15}$
$c_{10}^{(j)}$	$a^6$, odd	$-\frac{256}{45}$	$-\frac{352}{45}$

The two uses of $\alpha$ in the literature must not be conflated. In Refs. [4, 12], $\alpha$ labels contributions associated with analytically continued digamma structure in a particular decomposition of the superextremal contact terms. The superextremal regime itself is defined by $q>1$, not by the value of this label. In Ref. [29], by contrast, $\alpha$ multiplies an additional contact-term deformation of a candidate Compton amplitude. There $\alpha_{\rm C}=0$ is the baseline amplitude, which agrees with the BHPT result through the orders tested, including the terms labelled by the dissipative sign $\eta$ so that setting $\alpha_{\rm C}\ne0$ adds a distinct contact deformation. Our horizon-exchange construction has no independent $\alpha_{\rm C}$ parameter: it determines the full continued near-zone response, including the digamma-pole contributions. Removing those poles would define a separate subtraction prescription, not the $\alpha=0$ member of a horizon-exchange family. The possible identification of the particular contact expression discussed in the text with the $\alpha_{\rm C}=1$ term of Ref. [29] requires a term-by-term comparison of the full covariant amplitudes in matched kinematic and normalization conventions; it is not established by the polar limit or by the present response relation. No identification of the Bautista bookkeeping label with $\alpha_{\rm C}$ is made here. In particular, our horizon-continuation label $\eta$ is a path label in the monodromy problem and is not either contact-term parameter.

Analytic derivation of the absorptive response.

Write Eq. (8) as $S(\mathcal{K})=F[1+H(\mathcal{K})]$ with $H(\mathcal{K})=iE\mathcal{K}+O(\mathcal{K}^2)$, $E\equiv e^{i\pi\nu}+e^{-i\pi\nu}\rho$. Since $F$ and $\nu$ are invariant, the exchange-odd part is $S^-=\frac{i}{2}FE(\mathcal{K}_+-\mathcal{K}^{(2)}_-)$ up to terms quadratic in $\mathcal{K}_+$, which are $O(\epsilon^9)$ here. For integer $s$ and $\nu=\ell+\delta$:

(a) $\rho=\sin\pi(\delta+i\epsilon)/\sin\pi(\delta-i\epsilon)$ and, exactly, $E=(-1)^\ell e^{\pi\epsilon}\sin2\pi\delta/\sin\pi(\delta-i\epsilon)=(-1)^\ell e^{\pi\epsilon}(2i\delta/\epsilon)[1-i\delta/\epsilon+O(\epsilon^2)]$;

(b) $E=-e^{2i\pi\eta\delta}\sin(\pi\delta-i\pi\tau)/\sin(\pi\delta+i\pi\tau)=1+2i\pi\delta(\eta+\coth\pi\tau)+O(\delta^2)$. This Taylor expansion is at fixed nonzero $\tau$ and is not uniform as $\tau\to0$. In the axisymmetric and Schwarzschild limits one must retain the exact expression for $\Xi$ until it is combined with the $\tau$-dependent factor in $\mathcal{K}_{+}$, and only then take the limit. In particular, $\lim_{\tau\to0}\tau(1+\coth\pi\tau)=1/\pi$; the final expression (S34) is therefore regular at $\varpi=0$, with $\varpi/(e^{2\pi\varpi}-1)\to1/(2\pi)$. For $s=-2$, $\ell=2$:

(c) Anomalous dimension. The $n=0$ equation of the MST recurrence, $\beta_0+\alpha_0R_1+\gamma_0L_{-1}=0$, has no $O(\epsilon)$ term because the spheroidal eigenvalue $E(c)=6-\frac{4m}{3}c+\left[\frac{10}{189}(9-m^2)-1\right]c^2+O(c^3)$ cancels the explicit $mq\epsilon$ terms of $\beta_0$. At $O(\epsilon^2)$, $\alpha_0R_1=-\frac{5}{378}\epsilon^2(9\kappa^2+m^2q^2)$, $\gamma_0L_{-1}=O(\epsilon^4)$, and the $q^2$ terms cancel against those of $E(c)$, leaving $\delta=\nu_2\epsilon^2+O(\epsilon^3)$ with $\nu_2=-107/210$ for all $q$ and $m$. Numerically $\nu=2-\frac{5}{210}\epsilon^2+\frac{mq}{12}\epsilon^3+O(\epsilon^4)$; $\nu_3$ does not enter Eq. (57): once the powers of $\delta$ cancel, $\delta$ appears only through $\delta/\epsilon=\nu_2\epsilon+O(\epsilon^2)$.

(d) Response. From § 1.4, with

In the series ratio, $a_1=-\gamma_1/\beta_1=\frac{5i}{18}\epsilon\kappa(3-i\tau)$, while $a_{-1}=O(\epsilon^3)$ because $\alpha_{-1}\propto(\delta+i\epsilon)(\delta-i\epsilon)$; the $n=1$ terms of $X_\nu$ and of $X_{-\nu-1}$ give $X_{-\nu-1}/X_\nu=1-\frac{5}{9}\epsilon\kappa\tau+O(\epsilon^2)=1+\frac{5mq}{9}\epsilon+O(\epsilon^2)$. Altogether

(e) Far zone. At $O(\epsilon)$, $Fe^{-2i\epsilon\ln\epsilon}$ does not depend on $q$; at $q=0$ it receives $2i\epsilon\ln2$ from $2^{2i\epsilon}$, $i(2\gamma_E-\frac{25}{12})\epsilon$ from $\Gamma(1-i\epsilon)/\Gamma(5+i\epsilon)$ and $-i\epsilon$ from the $n=1$ term of the far-zone series, i.e.

Combining,

which with

and $\pi(\tau+\epsilon)=(\omega-m\Omega_H)/2T_H$ gives the leading exchange-odd amplitude, including its first relative correction:

This form makes the thermal factor and the first subleading MST and far-zone corrections explicit. Equivalently, the relative $O(\epsilon)$ correction can be generated by the following operator acting on the leading $\tau$-dependent factor (with $\partial_\tau$ differentiating that factor):

whose constant and $q$-dependent terms are independently tested in Sec. S7. Acting on the non-polynomial part of Eq. (47), the same operator also generates that of the order-$\epsilon^7$ coefficient in the exchange-even sector; there this is a numerical result (Sec. S7).

(f) Exchange-even part. There $(1+\Xi)/2=1+i\pi\delta(\eta+\coth\pi\tau)+O(\delta^2)$ multiplies $\mathcal{K}_{+}=(\delta^2\mathcal{K}_{+})\delta^{-2}[1+\delta L+O(\delta^2)]$, with $L=2\ln(2\epsilon\kappa)+\psi(3-i\tau)+\psi(-2-i\tau)+\mathrm{const}$. The finite part is proportional to $L+i\pi(\eta+\coth\pi\tau)$ of $2\ln(2\epsilon\kappa)+2\operatorname{Re}\psi(3+i\tau)+i\pi\eta+\mathrm{const}$, with the normalization of Eq. (15). The pole term $(i/900)\epsilon^4\Pi_m/\nu_2$ of $H^{+}$ (checked numerically) is absent from $S$, whose $\epsilon^4$ coefficient is a cubic polynomial in $q$: it cancels against $F$.

Numerical validation

The exchange law Eq. (6) agrees with MST to 60 digits for $s=0,-1,-2$, $\ell\leq3$, $q\in[0,0.9]\cup[1.5,3]$ and $M\omega\leq0.3$. Direct integration of the monodromy loops ($s=-2$, $\ell=m=2$, $q=0.6$, $\epsilon=0.3$) verifies the two factorizations of $M_{\infty}$, the Floquet eigenvalues and the lasso images to $10^{-8}$. The invariance of $\nu$ and of the far-zone factor holds to $10^{-57}$ and $10^{-51}$.

For the low-frequency expansion, $B_{6}=i\Pi_{m}/450$ is verified to $10^{-10}$; Eq. (13) is exact for all $m$ and reproduces the $q^{6}$ coefficients of Ref. [7]. Also $[q^{5}]C=16b_{2}$ to 20 digits and $C/q^{5}\to64/15$. The super-extremal $a^{5}$ coefficients agree with direct MST to $10^{-9}$. The operator in Eq. (58) is verified in both sectors: its constant terms agree to $3\times10^{-8}$, the coefficient $5m/9$ to $2\times10^{-8}$, and no $q^{2}$ term is found above $10^{-9}$.

The leading-small-$M\omega$ coefficients from Eqs. (27) and (30) agree to $5\times10^{-6}$ for $s=-2$, $\ell=2$ and all $m$, and to $2\times10^{-4}$ for the additional $\ell=3$ and $s=0$ cases. The $m=2$ super-extremal $a^{6}$ coefficients extrapolated over $300\leq q\leq5000$ are 1.89648157 and 1.89629630, versus direct MST values 1.89648161 and $256/135$. The detailed expansion of § 1.8 agrees with full MST for $\epsilon\leq4\times10^{-3}$: the $\mathcal{X}$-ratio coefficient to $10^{-6}$, the $O(\epsilon)$ coefficient of $E$ to $10^{-5}$, and the spheroidal coefficients $E_{1},E_{2}$ to $10^{-9}$.

The unit-flux normalization (Sec. S8) was tested for all $m$, $0\leq q\leq0.95$, $0.01\leq\epsilon\leq1.2$, including superradiance and the complex-$\nu$ branch. The Floquet-amplitude relations hold to $10^{-59}$, the factorized and direct $\mathcal{S}_{P}$ agree to $10^{-60}$, and $|\mathcal{S}_{P}|^{2}+P_{H}=1$ to $2\times10^{-59}$. The exchange-odd absorption formula holds for both $\eta$ to $2\times10^{-6}$; its fraction is $1/2$ at $q=0$ and $\kappa/(1+\kappa)$ for $m=0$, to $5\times10^{-7}$.

Flux balance and unit-flux normalization

For real-frequency gravitational perturbations we now fix the normalization needed to turn the conserved radial identity into an energy flux. We use the standard Kinnersley-tetrad, spin-$s=-2$ convention

The physical ingoing solution has the future-horizon behavior

Here $Z^{H}_{\ell m\omega}$ is the coefficient of the Kinnersley radial function itself; it is not the regular-horizon Weyl scalar or the horizon shear. The latter quantities are obtained from it by the Teukolsky–Starobinsky identities and the Kinnersley-to-Hawking–Hartle tetrad transformation.

Let $\kappa_{H}=2\pi T_{H}$ denote the (positive) surface gravity, to distinguish it from the dimensionless $\kappa=\sqrt{1-q^{2}}$ used in the main text, and set $D_{+}=r_{+}^{2}+a^{2}=2Mr_{+}$. In the separation-constant convention of the radial equation in the main text,

the squared gravitational Starobinsky constant is

The horizon-flux formula following from the shear/area balance and the Teukolsky–Starobinsky conversion is then

This is the standard Teukolsky horizon-absorption formula [11, 24], integrated over the full sphere in $G=c=1$ units with the Fourier normalization of Eq. (59). The factor $1/\sqrt{2\pi}$ gives unit full-sphere norm, so no additional $2\pi$ occurs. Restoring $G$ at fixed metric perturbation multiplies the flux by $1/G$. The last two factors are the squared moduli of the horizon-frequency factors $k_{H}+i\kappa_{H}$ and $k_{H}+2i\kappa_{H}$ generated by the four radial Starobinsky derivatives; equivalently, they are $(k_{H}^{2}+4\epsilon^{2})(k_{H}^{2}+16\epsilon^{2})$ for $\epsilon=\kappa_{H}/2$.

Thus, with $A_H := Z^H_{\ell m\omega}$, the normalization in Eq. (65) is explicitly

The sign of $\mathcal{F}_H$ is therefore exactly the sign of $k_H$; it vanishes at the superradiant threshold $\omega=m\Omega_H$.

The conservation law is the Teukolsky–Starobinsky cross-spin bilinear: the spin-$+2$ and spin-$-2$ solutions are related by fourth-order Starobinsky operators, and the bilinear is radius-independent. A same-spin Schrödinger current is not appropriate because the separated Teukolsky potential is generally complex. Evaluation at the horizon and infinity gives

with, in the same normalization, $\mathcal{F}^{\infty}_{\mathrm{in}}=128\omega^6\lvert B^{\mathrm{inc}}_{-2,\ell m\omega}\rvert^2/\lvert C_{\ell m\omega}\rvert^2$ and $\mathcal{F}^{\infty}_{\mathrm{out}}=\lvert B^{\mathrm{ref}}_{-2,\ell m\omega}\rvert^2/(2\omega^2)$, where $R_{\ell m\omega}\to B^{\mathrm{inc}}_{-2,\ell m\omega}r^{-1}e^{-i\omega r_*}+B^{\mathrm{ref}}_{-2,\ell m\omega}r^3e^{i\omega r_*}$ at large $r$. Hence

For a unit-flux-normalized physical reflection coefficient $S_{\ell m\omega}$, this is equivalently

For the Kinnersley radial normalization above, the energy reflection coefficient is

This is the Teukolsky–Press flux conversion; the raw radial ratio is not itself a unit-flux reflection amplitude. Equivalently, in the scattering normalization used in the Kerr tidal-response literature,

The MST factor $F_{\ell m}$ in Eq. (8) must consequently be understood as the Starobinsky-normalized ratio of the incoming and outgoing asymptotic coefficients of the $R^C_\nu$ Floquet solution:

The coefficients in the last ratio are defined by expanding that same Coulomb-wave MST solution at infinity in the $r^{-1}e^{-i\omega r_*}$ and $r^3e^{i\omega r_*}$ branches. With this definition, the response fraction in Eq. (8) is the remaining $\mathcal{K}$-dependent factor in the unit-flux reflection coefficient. Here $C_{\ell m\omega}=D+12iP M\omega$ is the complex Starobinsky constant for parity $P=\pm1$, with $D$ real and $\lvert C_{\ell m\omega}\rvert^2$ given by Eq. (62). The helicity-preserving partial-wave coefficient, which enters the Compton amplitude and was used for the expansions of Sec. S4, is the parity average and carries $D$ in place of $C_{\ell m\omega}$. It is not unit-flux normalized: $\lvert S_P\rvert^2=\lvert S\rvert^2+\lvert S_{\mathrm{flip}}\rvert^2$, with the helicity-reversing coefficient $S_{\mathrm{flip}}=(12iM\omega/D)S$, so $1-\lvert S\rvert^2-\mathcal{P}_H\simeq\epsilon^2/16$. Because $C_{\ell m\omega}/D=1+iP\epsilon/4+O(\epsilon^2)$ is exchange-even, Eqs. (13) and (14), the transcendental parts of Eqs. (47) and (15), the fixed-$a\omega$, small-$M\omega$ limits and all quoted coefficients are the same in both normalizations; only the coefficients of degree below five of $p^{(m)}$ and the $O(\epsilon)$ phase constant in Eqs. (57) and (S35), which shifts by $P/4$, refer to the helicity-preserving normalization.

Equation (S47) is checked by evaluating the two Floquet amplitudes independently from the same continued-fraction coefficients. The factorized $S$ agrees with $(C/16\omega^4)B^{\mathrm{ref}}/B^{\mathrm{inc}}$ and satisfies $\lvert S\rvert^2+\mathcal{P}_H=1$ to 59 digits.

The decomposition $S_\pm=S^{\mathrm{even}}\pm S^{\mathrm{odd}}$ is therefore useful for the physical outer-horizon solution. Once the unit-flux normalization has been made, its contribution to the absorption probability is

This is the precise amplitude-level sense in which the exchange-odd term contributes to the physical flux balance. It is not a difference of two horizon fluxes: the inner-horizon image is an auxiliary analytic solution, not a second physical exterior scattering state. In particular, the exchange-even part can also contribute to absorption, including the leading superradiant flux. With the unit-flux $F_{\ell m}$, the exchange-odd term of Eq. (71) agrees with Eq. (14) for both $\eta$ (with $\bar n\to\bar n+1$ for $\eta=-1$); it is one half of $\mathcal{P}_H$ at $q=0$, and the fraction $\kappa/(1+\kappa)$ of it for $m=0$ (Sec. S7).