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The minimum dissipation theorem fixes the boundary velocities and the far-field velocity: among admissible incompressible velocity fields, the Stokes flow minimizesFor a competitor $\mathbf{u}+\mathbf{v}$ with zero boundary data for $\mathbf{v}$, integration by parts eliminates the cross term and leaves $\mathcal{D}[\mathbf{u}+\mathbf{v}]-\mathcal{D}[\mathbf{u}]=2\mu \int \mathbf{e}(\mathbf{v}):\mathbf{e}(\mathbf{v})dV\ge 0$.

Extend the actual two-sphere velocity rigidly through sphere 2. It is an admissible field in the one-sphere exterior, with the same translation and rotation of sphere 1. It is continuous across the filled boundary, incompressible, and adds zero strain dissipation inside. Applying the theorem to the exact isolated-sphere flow therefore givesWith all applied couples and the second force zero, the boundary-work identity is ${\mathcal{D}}_{\mathrm{pair}}={\mathbf{F}}_1\cdot {\mathbf{U}}_1$. Thus $6\pi \mu a|{\mathbf{U}}_1{|}^2\le {\mathbf{F}}_1\cdot {\mathbf{U}}_1\le |{\mathbf{F}}_1||{\mathbf{U}}_1|$, which provesThis fixed-force comparison of minimum viscous dissipation uses a comparison at fixed actual velocity first; directly comparing different-force or different-velocity solutions would not justify the result.

If ${\mathbf{G}}_1\ne 0$, the power is instead ${\mathbf{F}}_1\cdot {\mathbf{U}}_1+{\mathbf{G}}_1\cdot {\mathbf{\Omega }}_1$, so the preceding estimate no longer bounds the force contribution alone. The two-sphere hydrodynamic mobility matrix generally has a nonzero self translation-rotation coupling: the freely moving second sphere reflects the first sphere's torque field. In the planar geometry this coupling produces a velocity perpendicular to $\mathbf{R}$; by choosing the sign and magnitude of the couple when the force has a component in that direction, its contribution to ${\mathbf{F}}_1\cdot {\mathbf{U}}_1$ can exceed the isolated force-only value. There is no universal force-only inequality with an additional applied couple. Special geometries can eliminate the coupling, but do not restore a general theorem.

Put ${\mathbf{W}}_2={\mathbf{F}}_2/\zeta$. To the needed accuracy the Faxén translation law gives ${\mathbf{U}}_2={\mathbf{W}}_2+\mathsf{M}\mathbf{V}+O({ϵ}^3|\mathbf{V}|)$. Enforcing its prescribed velocity therefore givesHere the force on sphere 1 remains prescribed; no extra second-order inversion term is needed in ${\mathbf{F}}_2$. Substitute this force into the incident field at sphere 1:The negative second-order term reflects the reduced force required on sphere 2 because sphere 1 already entrains it.

Sphere 2 is now force-free and torque-free. Its translation and spin follow the incident rigid-motion components, so it emits neither a Stokeslet nor a rotlet. Its first disturbance is a stresslet from the incident rate of strain, of magnitude $\mu a^3(a|\mathbf{V}|/R^2)$. Its return velocity at sphere 1 is consequently $O(a^4|\mathbf{V}|/R^4)$:More precisely, that leading correction is $-(15/4){ϵ}^4(\mathbf{V}\cdot \mathbf{n})\mathbf{n}$. It vanishes for purely transverse forcing, for which the first translational correction is of order ${ϵ}^6$.

The first-return spin requires an additional cancellation. The leading incident strain at sphere 2 isIt is axisymmetric about the line of centers, so the reflected stresslet has zero vorticity at sphere 1 on that axis. A naive vorticity estimate of order $(|\mathbf{V}|/a){ϵ}^5$ therefore has zero coefficient. The finite-size correction to the incident strain and the next force-free scattering multipoles are two powers of $ϵ$ smaller; these generically giveThis suppressed spin from a force-free distant sphere is also consistent with the reciprocal torque-to-self-translation coupling. Purely longitudinal forcing gives zero spin exactly by axial symmetry. Distinguishing the stresslet's on-axis curl from its generic off-axis scale is essential here.

For zero liquid flux the reduced gradients are ${\Gamma }_X=-4J/(H\Gamma )$ and $H_X=6J/(H^2\Gamma )$. Their ratio givesIntegrate $dX/dH=H^2\Gamma /(6J)$ to obtain the requested implicit height profile:The endpoint relation is $H_1^2=H_0^2+3\delta$, and the same expression at $X=1$ fixes $J$. For $H_0=0$ it simplifies tosoThis zero-liquid-flux surfactant film grows to the right from a formal dry edge, with $H\sim (18JX{)}^{1/3}$ there. Its dimensionless velocity isThe lower two thirds flow left, the upper third flows right, and the integrated liquid flux is zero. Surface advection carries the positive surfactant flux, with $U_s=H^2{H}_X/6=J/\Gamma$.

The flux curve obeys $dJ/d\delta =(\sqrt{3}/4)\sqrt{\delta }(1-\delta )$. It increases throughout the interval and reaches its formal maximum $J_{\max }=1/(5\sqrt{3})$ at $\delta =1$. Larger depletion strengthens the Marangoni driving, so increasing transport is plausible. But finite flux at a zero-concentration endpoint is a singular prediction: $U_s=J/\Gamma$ and the height gradient diverge there. Neglected diffusion, capillarity or an endpoint region must regularize the physical limit, and can alter its maximum. The dry initial edge is likewise outside a uniform small-slope approximation.

The profiles and velocity arrows illustrate the opposing interior-flow directions and the formal endpoint maximum; neither sketch treats the singular edges as resolved lubrication regions.

With zero surfactant flux and positive concentration, the nondiffusive surface velocity is zero. The reduced equations become ${\Gamma }_X=6Q/H^2$, $H_X=-12Q/H^3$. Integrating, with the prescribed initial height and concentration, givesAt the far end, the imposed concentration drop fixes $H(1{)}^2=H_0^2+4\delta$, henceThis steady surfactant film with zero surface flux rises to the right while its concentration decreases. The liquid flux is negative, despite the rightward Marangoni traction: the adverse hydrostatic gradient is strong enough to immobilize the surface and drive the interior leftwards.

In dimensionless height $Y=y/h_{\ast }$, the velocity profile isIt vanishes at both boundaries and is negative in the interior when $\delta >0$. Its integral is $Q=-H^3{H}_X/12$. For $\delta =0$, the nondegenerate film is flat and stationary. If its initial thickness is zero, the formal $X^{1/4}$ edge has an unbounded slope, so the lubrication approximation applies away from that edge rather than at the exact dry point.

Take $z$ positive upward, and write $b=\tilde{\rho }g$. For a straight steady viscous buoyant conduit, the inner and outer axial equations areRegularity gives $w_i^{\prime }(0)=0$, the wall has $w_o(R)=0$, and velocity and tangential traction are continuous at $r=a_0$: $w_i=w_o$, $\lambda \mu w_i^{\prime }=\mu w_o^{\prime }$. With no interfacial tension specified, normal traction is continuous; for this straight steady flow that sets $P_i=P_o$ at the interface. Thus the modified gradients are common in this case.

When their value is zero, integration givesConsequentlyThe large center-to-interface speed ratio explains why the inner parabolic component dominates at small viscosity ratio.

For a slowly varying radius, set $B=b-P_{i,z}$. The local inner solution has $w_i=w_a+B(a^2-r^2)/(4\lambda \mu )$, soThe shear-driven outer interface speed has scale $a^{2}B\ln (R/a)/\mu$. The pressure-driven return flow needed for zero total flux adds a velocity of order $q/R^2$ and a shear-driven contribution of order $a^{2}B/\mu$. Relative to the inner parabolic velocity these terms are $O(\lambda \ln (R/a))$ and $O((a/R{)}^2)$, which are small under the stated assumptions. Axial viscous derivatives are also smaller because the plume is slender. ThereforeFor clarity, the outer pressure-gradient estimate must include both return flux and transmitted interface shear. Their flux scales are $R^4{P}_{o,z}/\mu$, $a^2{R}^{2}B/\mu$ and $q$. Zero total flux hence impliesThis justifies the negligible outer pressure gradient. We subsequently choose its negligible pressure reference as zero, as in the reduced model.

Near the moving interface the outer radial source flow is $u_r=a{a}_t/r$: the interface's axial advection is smaller because its axial velocity is small compared with the inner core speed. Its radial strain is ${\partial }_r{u}_r=-a{a}_t/r^2$. Continuity of normal traction at $r=a$, neglecting the much smaller inner normal viscous stress, givesWith $A=a^2/a_0^2$, this is $P_i=\mu A_t/A$ after dropping $P_o$. Conservation of plume volume is $\pi (a^2{)}_t+q_z=0$.

Let the axial and velocity scales beSince $\mu U_0/(b{\ell }^2)=1$, substitution gives the conduit equationThe buoyant flux and the viscous normal-stress pressure are both retained at this scaling.

Linearizing about unit area gives $a_T+2a_Z-a_{TZZ}=0$. ThusWave crests propagate upward. The group velocity is $2(1-k^2)/(1+k^2{)}^2$: it is downward for $|k|>1$, so the phase direction is not the direction of every wave packet.

For a travelling wave $f(\zeta )$, integrating once and using its uniform far field givesDivide by $f^2$, multiply by $f^{\prime }/f$ and integrate again. A convenient potential, defined up to an additive constant, isAt the positive crest $f=\alpha f_0$, the derivative is zero. Subtracting the far-field potential yieldsThis is the solitary-wave amplitude-speed relation for the conduit equation. A nontrivial elevation wave has $\alpha >1$ and $c>2f_0$; as $\alpha \downarrow 1$, the speed tends to the long-wave speed $2f_0$. Positive area is required throughout the derivation.

Let the actual dimensional thickness be $d=ϵah$, and let $\bar{z}$ increase downward. The leading curvature of the outer free surface isThe Young–Laplace equation gives $p_{\bar{z}}=-\gamma (d_{\bar{z}}/a^2+d_{\bar{z}\bar{z}\bar{z}})$. In the lubrication approximation, the downward velocity at distance $y$ from the wall is $(\rho g-p_{\bar{z}})(dy-y^2/2)/\mu$, from no slip and zero free-surface shear. Integrating this profile givesCircumferential corrections are higher order in $d/a$, so volume conservation is $d_{\bar{t}}+q_{\bar{z}}=0$ at leading order. DefineThen the thin liquid film on a vertical cylinder satisfiesThe circumferential-curvature term $h_z$ is retained along with axial curvature.

The perturbation calculation givesLong waves with $0<|k|<1$ grow, while $|k|>1$ decay; the fastest growth is $1/12$ at $|k|=1/\sqrt{2}$. Disturbances drift downward at speed $G$. The dimensional unstable wavelengths exceed $2\pi a$, with most unstable wavelength $2\pi \sqrt{2}a$. This is the thin-film version of the Rayleigh–Plateau instability driven by circumferential curvature and opposed by short-wave axial curvature.

For a travelling wave, one integration fixed by $H\to 1$ givesor $H^{\prime \prime \prime }+H^{\prime }=3c(H-1)/H^3-G(1-H^{-3})$. In the large positive-speed core, substitute the expansion using the supplied assumptions. The left side at order $c^{2/3}$ is $H_2^{\prime \prime \prime }+H_2^{\prime }$, while the gravity term is order one and $3c/H^2$ is only order $c^{-1/3}$. Thus $H_2^{\prime \prime \prime }+H_2^{\prime }=0$. Its general solution is a constant plus a sine and cosine. The vanishing leading height and slope at both endpoints selectThe core length is $2\pi$ in the axial coordinate; its amplitude is fixed by transition matching.

At either edge set $X=(3c{)}^{1/3}(x-x_e)$, with $x_e=0$ at the trailing edge and $x_e=2\pi$ at the leading edge. Keeping the height $H$ of order one, the exact wave equation becomesThus the transition equation is $H^3{H}_{XXX}=H-1$. Its linearization about $H=1$ has exponents $1$ and $-1/2\pm i\sqrt{3}/2$. Behind the pulse, approaching the uniform film as $X\to -\infty$ permits only the positive real exponent. The positive-growing, nontrivial branch has one amplitude that is removed by translation, giving the relevant unique $H_{-}$. In front, decay as $X\to +\infty$ permits the two real oscillatory amplitudes; translation removes one parameter and leaves a one-parameter family $H_{+}$. These are the unstable manifold and stable manifold dimensions for the transition system. The literal boundary condition also admits the constant solution and an opposite departing branch, so uniqueness here is specifically for the branch matching a raised positive pulse.

Match the supplied trailing quadratic asymptotic to the core near zero. It gives $K_2=\kappa 3^{2/3}$, because $\frac{1}{2}{c}^{2/3}{K}_2{x}^2=\frac{1}{2}\kappa (3c{)}^{2/3}{x}^2$. The leading maximum occurs at $x=\pi$, soThe front tail has the form $H-1\sim C{e}^{-X/2}\cos (\sqrt{3}X/2+\theta )$. Its capillary waves occur ahead of the descending pulse, not behind it; the trailing tail is monotone to leading order. Converting its frequency to physical distance gives

This large solitary pulse on a cylindrical film sketch shows the leading core and the front-tail eigenmodes; the edge transitions are schematic and the ringing is enlarged for visibility. The thin-film asymptotics require both $c\gg 1$ and $ϵc^{2/3}\ll 1$, rather than taking speed to infinity at fixed film slenderness.

At the next core order, $H_0^{\prime \prime \prime }+H_0^{\prime }=-G$. Integrating and imposing zero $H_0^{\prime }$ at both endpoints givesThe constants in the two quadratic transition asymptotics match the core endpoint heights: $A_0+B_0={\beta }_{-}$ and $A_0+B_0-2\pi G={\beta }_{+}$. Hence $2\pi G={\beta }_{-}-{\beta }_{+}$, and the dimensional uniform thickness selected by matching isWith the supplied constants this fixes $G=3.75/(2\pi )$; $B_0$ is not separately fixed by these height conditions. The thickness selection and capillary wavelength follow from matched asymptotic expansion, not from the linear instability calculation.

Introduce the slow time $T=\epsilon t$ and treat $t,T$ independently in the method of multiple scales. WriteAt the next order,For fixed $R$ define the period average byHere $f$ inside the averages is evaluated at the leading position and velocity $R\cos \varphi ,-R\sin \varphi$. Orthogonality to both fundamental harmonics is the solvability condition in the method of multiple scales: it removes the resonant forcing that would otherwise generate a secular term. Since the squared sine and cosine averages are $1/2$, the amplitude-phase equations areEquivalently $\dot{R}=-\epsilon ⟨f\sin (t+\theta )⟩$ and $\dot{\theta }=-\epsilon ⟨f\cos (t+\theta )⟩/R$ at first order. These equations describe a bounded, weakly perturbed oscillation over $t=O({\epsilon }^{-1})$ while the amplitude remains in a range where the expansion is ordered. At zero amplitude the oscillation phase coordinate is singular; Cartesian harmonic coefficients or an equilibrium analysis should replace it.

Keep the prescribed outer slip velocity fixed when perturbing the unsteady Prandtl equation. Cancelling the base equation and retaining terms linear in the disturbance gives the linearized unsteady Prandtl equationThe no-slip boundary condition supplies $\tilde{u}=\tilde{v}=0$ at the wall; the fixed outer velocity supplies $\tilde{u}\to 0$ at infinity. There is no prescribed zero normal-velocity disturbance at infinity: a finite displacement-related value is permitted.

With the frozen parallel profile $U=U(y)$, $V=0$, the coefficients are invariant under translations in $x$ and $t$. Normal modes therefore separate those variables. For the mode convention used here the equations areEliminate $u=-v^{\prime }/(i\alpha )$ to obtain the Prandtl normal-mode equationwith a bounded far-field $v$. Requiring $v(\infty )=0$ would incorrectly eliminate the proposed outer profile. For real $U$ and real $\alpha$, complex conjugation of the equation and its boundary conditions replaces $(\alpha ,c,v)$ by $(-\alpha ,c^{\ast },v^{\ast })$. This proves the stated eigenvalue symmetry. The paired modes have the same temporal growth rate $\alpha \mathrm{Im}c$, so choosing positive wavenumber loses no real physical disturbance.

Here the prescribed parallel profile is a local frozen-coefficient model. A nontrivial arbitrary $U(y)$ with $V=0$ and constant outer slip is not generally an exact steady solution of the unforced unsteady Prandtl equation; its base equation would require $U^{\prime \prime }=0$. The following stability calculation uses the simplifying local model stipulated for the mode analysis.

Set $c_0=U_c=U(y_c)$. Away from the critical point the dominant equation is $(U-c_0)v_0^{\prime }-U^{\prime }{v}_0=0$, whose solutions are multiples of $U-c_0$. Choosing the coefficient to be zero below $y_c$ and one above it satisfies the wall and tangential far-field conditions. The choice $c_0=U_c$ makes $v_0$ continuous across the joining point; $U^{\prime }(y_c)=0$ also makes its first derivative continuous. Its second derivative jumps. Thus it is a valid leading outer solution away from $y_c$, but viscosity must smooth the join in a critical layer in a shear flow.

Near the join, $U-c_0\sim (y-y_c{)}^2/2$. If the layer width is $\ell$, matching gives $v=O({\ell }^2)$. The advection side of the mode equation scales as $\alpha {\ell }^3$ and $v^{\prime \prime \prime }$ scales as ${\ell }^{-1}$. Balance gives ${\ell }^4\sim {\alpha }^{-1}=\epsilon /k$. Equivalently, the eigenvalue correction ${\epsilon }^{1/2}{c}_1$ balances $(y-y_c{)}^2$, so the quarter-power critical layer hasSubstitution, retaining the leading terms and cancelling their common factor, gives

At order ${\epsilon }^{1/2}$ outside the layer,A solution consistent with the chosen lower branch and the wall conditions is $v_1=0$ below the join. Above it a particular solution is $v_1=-c_1$; a multiple of $U-c_0$ represents an arbitrary amplitude renormalization and may be set to zero. Matching consequently requiresThe corresponding derivatives match as $w_Y\sim Y$, $w_{YY}\to 1$ on the positive side and as zero on the negative side. The matching statements require that growing homogeneous corrections be absent.

To symmetrize these different end conditions and remove $k$ and $i$, chooseThe polynomial $A(z)=z^2/2-C$ is itself a solution of the transformed homogeneous equation. Direct substitution, using $a^4=ik$, givesThe derivative is third order, as required by the original mode equation. Initially the two ends are along the rotated rays $z=aY$ with real $Y$. Continuing them to the real $z$ axis is compatible with the asymptotic end conditions: the decreasing homogeneous correction behaves exponentially as $\exp [-z^2/(2\sqrt{2})]$ and still decreases throughout the rotation from angle $\pi /8$ to zero. Thus the supplied real-axis spectral normalization uses the same recessive conditions; a complex coordinate change should not be mistaken for a real stretching alone.

For each supplied real eigenvalue $C_n=(4n+7)/\sqrt{2}$, the phase-speed correction isThe mode factor has temporal magnitude $e^{\alpha \mathrm{Im}ct}$. ThereforeAmong the listed indices, $n=-2$ gives positive growth $\gamma \sim \sqrt{\alpha }/2$; $n=-1,1,2,\dots$ give decay. Hence the frozen non-monotone flow has high-wavenumber instability of a non-monotone Prandtl layer, with arbitrarily rapid linear growth as the positive wavenumber increases. The normal-mode calculation demonstrates linear instability; it is not by itself a claim about the final nonlinear state.