2.1. The inviscid and viscous RP modes ($h=0$)

The classical RP modes are unstable axisymmetric Fourier modes satisfying $0 < kR_0 < 1$ for the unforced system ($h=0$). These are governed by the following inviscid ((2.2a), Rayleigh Reference Rayleigh1878) and viscous dispersion relation (Rayleigh Reference Rayleigh1892a; Weber Reference Weber1931; Chandrasekhar Reference Chandrasekhar1981; Liu & Liu Reference Liu and Liu2006) with growth rate $\sigma _0$ (inviscid) and $\sigma$ (viscous), respectively:

(2.2a)\begin{gather} \sigma_0^2 = \frac{T}{\rho R_0^3}kR_0(1 - k^2R_0^2)\frac{{\mathrm{I}}_1(kR_0)}{{\mathrm{I}}_0(kR_0)}, \end{gather}

(2.2b)\begin{gather} \left. \begin{gathered} \sigma^{2}+2 v k^{2}\left[\frac{{\mathrm{I}}_{1}^{\prime}(k R_{0})}{{\mathrm{I}}_{0}(k R_{0})}-\frac{2 k l}{l^{2}+k^{2}} \frac{{\mathrm{I}}_{1}(k R_{0})}{{\mathrm{I}}_{0}(k R_{0})} \frac{{\mathrm{I}}_{1}^{\prime}(l R_{0})}{{\mathrm{I}}_{1}(l R_{0})}\right]\sigma - \left(\frac{l^{2}-k^{2}}{l^{2}+k^{2}}\right)\sigma_0^2 = 0,\\ \text{where}\quad l^2 \equiv k^2 + \frac{\sigma}{\nu}. \end{gathered} \right\} \end{gather}

Here ${\mathrm {I}}_{m}(z)$ is the $m$th order modified Bessel function of the first kind and ${\mathrm {I}}_{m}^{'}(z) \equiv {\textrm {d}{\mathrm {I}}_m}/{\textrm {d}z}$. In figure 2, $\sigma _0$ and $\sigma$ are obtained by numerically solving (2.2a) and (2.2b) for the inviscid and viscous cases, respectively. Unlike the inviscid relation (2.2a) which is quadratic in $\sigma _0$, the viscous dispersion relation given by (2.2b) is transcendental in $\sigma$. It admits in addition to two capillary modes, a countably infinite set of hydrodynamic (or vorticity) modes as its roots and the latter are purely damped modes (García & González Reference García and González2008). In figure 2 we only depict the growth and decay rates corresponding to the two capillary modes in the range $0 < kR_0 < 1$ for different values of Ohnesorge number ${Oh} = {\mu }/{\sqrt {T\rho R_0}}$. Our aim in this study is to stabilise the capillary modes in the range $0 < kR_0 < 1$ using radial forcing and this is discussed below.

Figure 2.

Inviscid and viscous growth (and decay) rates of RP modes ($0 < kR_0 < 1$) from numerically solving (2.2a) and (2.2b) (Weber Reference Weber1931; García & González Reference García and González2008). At any Ohnesorge ($Oh$) and $k$ in the range $0 < k < R_0^{-1}$, there are two capillary modes, one unstable ($\sigma > 0$) and another stable $(\sigma < 0)$. We stabilise the exponentially growing mode by forcing at $\varOmega \gg \sigma _{max}$, where $\sigma _{max}$ is the growth rate of the fastest growing RP mode, it being highest for the inviscid case ($Oh=0$) for $(kR_0)_{max} \approx 0.69$ with $\sigma _{{max}} \approx 0.34\sqrt {{T}/{\rho R_0^3}}$.

2.2. Dynamic stabilisation of RP modes – linear inviscid theory

The inviscid results on RP stabilisation using radial forcing were presented earlier in Patankar et al. (Reference Patankar, Farsoiya and Dasgupta2018) and are summarised very briefly here, for self-containedness. In the presence of radial forcing $\mathcal {F}(r,t) = -h({r}/{R_0})\cos (\varOmega t)$ and under the linearised, inviscid, irrotational approximation, the equation governing the amplitude $a_m(t;k)$ of standing waves on the free surface of the form $\eta (z,\theta,t) = a_m(t;k)\cos (m\theta )\cos (kz)$ is the Mathieu equation (2.3)

(2.3)\begin{equation} \frac{{\rm d}^2a_m}{{\rm d}t^2} + \frac{{\mathrm{I}}^{'}_{m}\left(kR_0\right)}{{\mathrm{I}}_{m}\left(kR_0\right)}\left[\frac{T}{\rho R_0^3}kR_0\left(k^2R_0^2 + m^2 -1\right) + kh\cos(\varOmega t)\right] a_m(t;k) = 0. \end{equation}

The stability diagram for (2.3) may be obtained using Floquet analysis (Patankar et al. Reference Patankar, Farsoiya and Dasgupta2018). For $h \neq 0$, we have the interesting prediction that axisymmetric unstable RP modes can be stabilised by choosing $h$ to be sufficiently large. This is readily seen in the stability chart in figure 3(a) where the solid curve in black indicates the threshold value of forcing $h$ above which, a RP mode is stable. The line in blue indicates all unstable RP modes for $h=0$. Two representative RP unstable modes are chosen, viz. $k_0=4.8$ cm$^{-1}$ (wavelength $\lambda \approx 1.309$ cm) and $k_0=3.48$ cm$^{-1}$ ($\lambda \approx 1.8$ cm) . The plot predicts the threshold values of forcing strength $h_{{cr}}=1.21\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$ and $h_{{cr}}=4.17\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$, respectively, beyond which these modes can be stabilised. For generating figure 3(a), we have chosen $\varOmega = 600 {\rm \pi}\,\textrm {rad}\,\textrm {s}^{-1}$ ($f=300$ Hz), $R_0=0.2$ cm, density $\rho =0.957\,\textrm {gm}\,\textrm {cm}^{-3}$, surface tension $T=20.7\,\textrm {dyn}\,\textrm {cm}^{-1}$. These fluid parameters approximately correspond to silicone oil (Vega & Montanero Reference Vega and Montanero2009) with its viscosity artificially set to zero. Note that at these forcing frequencies we may safely ignore compressibility effects as may be inferred from the order of magnitude of the two typical velocity scales, viz. $\textrm {maximum}[{h_{c}}/{f},fR_0] \approx 139\,\textrm {cm}\,\textrm {s}^{-1}$ for $f=300$ Hz and $h_c=4.17\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$. This is negligible compared with the typical acoustic speed ${O}(10^5)\,\textrm {cm}\,\textrm {s}^{-1}$ in the fluid at ambient conditions.

Figure 3.

Grey and white indicate unstable and stable regions, respectively. Panel (a) shows the inviscid stability chart for (2.3). The forcing frequency $f = 300\,\textrm {Hz} \gg \sigma _{{max}}= 0.34\sqrt {T/(\rho R_0^3)} = 17.68\,\textrm {Hz}$. Parameters are for case 1 in table 3 with $\mu ^{I}=0$. Panel (b) shows the time signal (red curve) from the numerical solution to the 3-D Euler equation (Popinet Reference Popinet2014) with an RP mode ($k_0=4.8\,\textrm {cm}^{-1},m_0=0$) excited at $t=0$. Black curve: solution to (2.3). Left inset: zoomed out view of solution to (2.3). Right inset: stability chart for $m=4$. An unstable non-axisymmetric Fourier mode ($k=28.8=6k_0,m=4$ in the grey region) at $\tilde {t} \approx 14$ s causes destabilisation of the cylinder. (a) Stability plot. (b) Result shown for $k=4.8,h= 1.8\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$.

The results from the numerical simulations discussed below are obtained using the open-source code Basilisk (Popinet Reference Popinet2014). The simulations are described in detail later on in § 5 and are briefly discussed here for consistency. The domain consists of a liquid cylinder of silicone oil surrounded by an ambient fluid of negligible density and viscosity (see figure 7 in § 5). The cylinder is subjected to a radial harmonic (in time) forcing. The simulation parameters for the case discussed below correspond to case $1$ from table 3 with $\mu ^{I} = \mu ^{O} = 0$ to reflect the inviscid limit that we analyse first. Figure 3(b) presents the time signal obtained from inviscid numerical simulations (Popinet Reference Popinet2014) for the axisymmetric mode $k_0=4.8,m_0=0$ excited at $t=0$. Note that this is a RP unstable mode and as seen from figure 3(a), it is expected to be stabilised beyond a threshold forcing of $h=1.21\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$. In figure 3(b) we see agreement between the solution to (2.3) and the numerical simulation for very brief time (approximately three forcing time periods), after which the signal from the numerical simulation begins to deviate and grow rapidly (around $\tilde {t}\approx 14$) in contrast to the solution to (2.3) which stays bounded (see left inset). A Fourier analysis of the interface at $\tilde {t} \equiv t\varOmega /2{\rm \pi} \approx 14$, indicated by the arrow, reveals the appearance of a non-axisymmetric mode $(k=28.8,m=4)$ in the simulation. This is a stable mode in the unforced system ($h=0$) but is destabilised at the imposed level of forcing, lying inside a tongue as seen in the right inset of figure 3(b). Thus, in the inviscid case the stabilisation that is achieved is only a quasi-stabilisation in the sense that while the RP modes can be rendered stable via forcing, concomitantly, other modes become unstable at the chosen level of forcing. It thus becomes clear that for obtaining dynamic stabilisation, we need to ensure that all Fourier modes either present initially in the system or born via nonlinear effects, both axisymmetric and three dimensional, should remain linearly stable at the imposed level of forcing. We will demonstrate in the next section that by taking viscosity into account and using the forcing frequency as a tuning parameter, this may be achieved.

2.3. Dynamic stabilisation of RP modes – linear viscous theory

Having demonstrated the inadequacy of dynamic stabilisation of RP modes in an inviscid model, we proceed to the viscous case. The motivation for including viscosity is simple to understand: it is known that inclusion of viscosity leads to displacement of the instability tongues upwards on the $h$–$k$ plane and these no longer touch the wavenumber axis (Kumar & Tuckerman Reference Kumar and Tuckerman1994). Our expectation is that by suitably choosing viscosity and the forcing frequency, we will be able to shift the unstable tongues sufficiently above the wavenumber ($k$) axis. This generates a sufficiently large stable region where not only the axisymmetric RP unstable mode ($k_0$) is stabilised (with forcing) but all higher modes accessible to the system are also stable. Note that the upward movement of the tongues occur not only for axisymmetric modes but also for non-axisymmetric ones. In particular, we will also see that for fixed viscosity, we can move the minima of the tongue upwards by increasing the forcing frequency. The algebra for the viscous analysis is somewhat lengthy and details are provided in the supplementary material available at https://doi.org/10.1017/jfm.2022.533. We outline the important steps that follow. Expressing all quantities as the sum of base plus perturbation, i.e.

(2.4a–c)\begin{equation} \hat{p} = p_b + p,\quad \hat{\boldsymbol{u}} = \boldsymbol{0} + \boldsymbol{u}\quad \text{and} \quad \text{perturbed free surface at}\ r = R_0 + \eta. \end{equation}

Substituting (2.4a,b) into the incompressible Navier–Stokes equations and linearising about the base state we obtain the equations governing the perturbations, viz.

(2.5a,b)\begin{equation} \left(\frac{\partial}{\partial t} - \nu \Delta\right)\boldsymbol{u} ={-}\frac{1}{\rho}\boldsymbol{\nabla}p,\quad \boldsymbol{\nabla}\boldsymbol{{\cdot}}\boldsymbol{u} = 0, \end{equation}

where the vector Laplacian of the incompressible velocity field is $\Delta \boldsymbol {u} \equiv -\boldsymbol {\nabla }\times \boldsymbol {\nabla }\times \boldsymbol {u}$. The linearised boundary conditions are obtained by substituting (2.4a–c) into the boundary conditions (supplementary material), employing Taylor expansion and retaining terms linear in the perturbation variables, viz. $\boldsymbol {u}, p$ and $\eta$ (the perturbation velocity $\boldsymbol {u}$ is written in terms of its components $(u_r,u_{\theta },u_z)$), we obtain

(2.6a)\begin{gather} \frac{\partial\eta}{\partial t} = u_r(r=R_0), \end{gather}

(2.6b,c)\begin{gather} \mu\left(\frac{\partial u_r}{\partial z} + \frac{\partial u_z}{\partial r}\right)_{r=R_0} = 0, \quad \mu\left(r\frac{\partial}{\partial r}\left(\frac{u_{\theta}}{r}\right) + \frac{1}{r}\frac{\partial u_r}{\partial \theta}\right)_{r=R_0} = 0, \end{gather}

(2.6d)\begin{gather} \left. \begin{aligned} & \left( \frac{\partial}{\partial r} + \frac{1}{r}\right)\left[\frac{\partial u_r}{\partial t} - \nu\left\{\Delta u_r - \frac{u_r}{r^2} - \frac{2}{r^2}\left(\frac{\partial u_{\theta}}{\partial\theta}\right)\right\}\right] +\mathcal{F}(r,t)\Delta_O\eta - 2\nu\Delta_O\left(\frac{\partial u_r}{\partial r}\right) \\ & \qquad ={-}\frac{T}{\rho R_0^2}\Delta_O\left[\eta + \left(\frac{\partial^2\eta}{\partial \theta^2}\right) + R_0^2\left(\frac{\partial^2\eta}{\partial z^2}\right)\right] \quad \text{at} r=R_0,\\ & \quad \text{with}\quad\Delta_{O} \equiv \frac{1}{r^2}\frac{\partial^2}{\partial \theta^2} + \frac{\partial^2}{\partial z^2}, \end{aligned} \right\} \end{gather}

(2.6e)\begin{gather} \boldsymbol{u}(r\rightarrow 0 ,t)\rightarrow \text{finite}, \end{gather}

where $\Delta$ is the scalar Laplacian in cylindrical coordinates. Equations (2.6a–e) are the linearised versions of the kinematic boundary condition (2.6a), the zero shear stress condition(s) at the free surface (2.6b,c), the normal stress condition at the free surface due to surface tension (2.6d) and the finiteness condition at the axis of the cylinder (2.6e), respectively. Equation (2.6d) has been obtained by eliminating pressure from the primitive form of the pressure jump boundary condition (see supplementary material). Note the presence of the forcing term $\mathcal {F}(r,t)$ in the normal stress boundary condition in (2.6d) indicating the time periodicity of the base state.

We solve (2.5a,b) in the streamfunction-vorticity formulation and, for this, the curl and double curl of (2.5a) leads to ($\boldsymbol {\omega } \equiv \boldsymbol {\nabla }\times \boldsymbol {u}$)

(2.7a,b)\begin{equation} \frac{\partial\boldsymbol{\omega}}{\partial t} = \nu\boldsymbol{\Delta}\boldsymbol{\omega},\quad \frac{\partial}{\partial t}\boldsymbol{\Delta} \boldsymbol{u} = \nu \boldsymbol{\Delta} \boldsymbol{\Delta} \boldsymbol{u}, \end{equation}

where $\boldsymbol {\Delta }$ is the vector Laplacian. Employing the toroidal-poloidal decomposition (Marqués Reference Marqués1990; Boronski & Tuckerman Reference Boronski and Tuckerman2007; Prosperetti Reference Prosperetti2011), the velocity and vorticity fields are expressed in terms of two scalar fields $\psi (r,\theta,z,t)$ and $\xi (r,\theta,z,t)$ using the decomposition

(2.8a,b)\begin{align} \boldsymbol{u} = \boldsymbol{\nabla}\times\left(\psi\hat{\boldsymbol{e}}_z\right) + \boldsymbol{\nabla}\times\boldsymbol{\nabla}\times\left(\xi\hat{\boldsymbol{e}}_z\right),\quad \boldsymbol{\omega} \equiv \boldsymbol{\nabla}\times\boldsymbol{\nabla}\times\left(\psi\hat{\boldsymbol{e}}_z\right) + \boldsymbol{\nabla}\times\boldsymbol{\nabla}\times\boldsymbol{\nabla}\times\left(\xi\hat{\boldsymbol{e}}_z\right), \end{align}

where $\hat {\boldsymbol {e}}_z$ is the unit vector along the axial direction of the cylinder (Boronski & Tuckerman Reference Boronski and Tuckerman2007). By construction the velocity field in (2.8a) is divergence free and it can be shown (see supplementary material) that the equations governing the toroidal and poloidal fields $\psi (r,z,\theta,t)$ and $\xi (r,z,\theta,t)$, respectively, are the fourth- and sixth-order equations

(2.9a,b)\begin{equation} \left(\frac{\partial}{\partial t} - \nu\Delta\right)\Delta_{{H}}\psi = 0 \quad \text{and}\quad \left(\frac{\partial}{\partial t} - \nu\Delta\right)\Delta\Delta_{{H}}\xi = 0, \end{equation}

where the scalar Laplacian

(2.10)\begin{equation} \Delta \equiv \frac{1}{r}\frac{\partial}{\partial r}(r\frac{\partial}{\partial r}) + \frac{1}{r^2}\frac{\partial^2}{\partial\theta^2} + \frac{\partial^2}{\partial z^2}= \Delta_{{H}} + \frac{\partial^2}{\partial z^2}. \end{equation}

As we have raised the order of our governing equations by taking curl and double curl, we need extra equations to determine the additional constants of integration. It was shown in Marqués (Reference Marqués1990) that this takes the form of an additional equation also known as the compatibility condition (Boronski & Tuckerman Reference Boronski and Tuckerman2007). For the present problem at linear order, this extra equation is simply the radial component of the vorticity equation (2.7a) (Boronski & Tuckerman Reference Boronski and Tuckerman2007), i.e.

(2.11)\begin{equation} \left. \begin{gathered} \dfrac{\partial \omega_r}{\partial t} = \nu\left\{\Delta \omega_r - \dfrac{\omega_r}{r^2} - \dfrac{2}{r^2}\left(\dfrac{\partial \omega_{\theta}}{\partial\theta}\right)\right\}, \\ \text{with}\quad \omega_r = \dfrac{\partial^2 \psi}{\partial r \partial z } -\dfrac{1}{r}\dfrac{\partial}{\partial\theta}\left(\Delta \xi \right) \quad \text{and} \quad \omega_\theta=\dfrac{1}{r}\dfrac{\partial^2 \psi}{\partial z \partial \theta } + \dfrac{\partial}{\partial r}\left(\Delta \xi\right). \end{gathered} \right\} \end{equation}

In order to determine the scalar fields $\psi (r,\theta,z,t), \xi (r,\theta,z,t)$, we need to solve (2.9a,b). Analogous to the inviscid analysis in Patankar et al. (Reference Patankar, Farsoiya and Dasgupta2018) we seek three-dimensional standing wave solutions of the form

(2.12a–c)$$\begin{gather} \psi(r,\theta,z,t) = \varPsi_m(r,t;k)\sin(m\theta)\cos(kz),\quad\xi(r,\theta,z,t) = \varXi_m(r,t;k)\cos(m\theta)\sin(kz), \nonumber\\ \eta(\theta,z,t) = a_m(t;k)\cos(m\theta)\cos(kz), \end{gather}$$

where $k \in \mathbb {R}^{+}$ and $m\in \mathbb {Z}^{+}$. Substituting (2.12a,b) into (2.9a,b) we obtain the equations governing $\varPsi _m(r,t;k)$ and $\varXi _m(r,t;k)$, viz.

(2.13a,b)\begin{equation} \left. \begin{aligned} & \left(\frac{\partial}{\partial t} - \nu \mathcal{L}\right)\mathcal{L_H}\varPsi_m = 0,\quad \left(\frac{\partial}{\partial t} - \nu \mathcal{L}\right)\mathcal{L}\mathcal{L_H}\varXi_m = 0,\\ & \text{where}\quad \mathcal{L_H} \equiv \frac{\partial^2}{\partial r^2} + \frac{1}{r}\frac{\partial}{\partial r} - \frac{m^2}{r^2} \quad \text{and} \quad \mathcal{L} \equiv \mathcal{L}_H - k^2. \end{aligned} \right\} \end{equation}

Our task now is to determine the linear stability of the (time-dependent) base state by identifying unstable and stable regions via Floquet analysis. This is indicated on the strength of forcing ($h$) vs wavenumber ($k,m$) plane for chosen fluid parameters $\rho,\nu, T$ and forcing frequency $\varOmega$ and is done in the next subsection.

2.3.1. Floquet analysis

Using the Floquet ansatz for time periodic base states, we assume the following forms for $\varPsi _m(r,t;k), \varXi _m(r,t;k)$ and $a_m(t;k)$ in (2.12a–c) (Kumar & Tuckerman Reference Kumar and Tuckerman1994)

(2.14a–c)\begin{equation} \left. \begin{aligned} & \varPsi_m(r,t;k) = \exp(\lambda_m(k) t)\sum_{n ={-}\infty}^{\infty}\tilde{\psi}_n^{(m)}(r;k)\exp({\rm i}\,n\varOmega t),\\ & \varXi_m(r,t;k) = \exp(\lambda_m(k) t)\sum_{n ={-}\infty}^{\infty}\tilde{\xi}_n^{(m)}(r;k)\exp({\rm i}\,n\varOmega t),\\ & a_m(t;k) = \exp(\lambda_m(k) t)\sum_{n ={-}\infty}^{\infty}\mathcal{M}_n\exp({\rm i}\,n\varOmega t), \end{aligned} \right\} \end{equation}

with $\lambda _m(k)$ being the Floquet exponent and $\tilde {\psi }_n^{(m)}(r;k)$ and $\tilde {\xi }_n^{(m)}(r;k)$ the complex eigenfunctions for each Fourier mode $(k,m)$. The complex eigenfunctions satisfy the reality condition $\tilde {\psi }_{-n}^{(m)} = (\tilde {\psi }_n^{(m)})^{*}$ and $\tilde {\xi }_{-n}^{(m)} = (\tilde {\xi }_n^{(m)})^{*}$, the superscript $^{*}$ indicating complex conjugation.

We substitute (2.14a,b) into (2.13a,b) respectively yielding fourth- and sixth-order differential equations (eigenvalue problems) governing $\tilde {\psi }_n^{(m)}(r;k)$ and $\tilde {\xi }_n^{(m)}(r;k)$ for each $n$ in the expansion (2.14a,b),

(2.15a)$$\begin{gather} \boldsymbol{O}^{(k,m)}\boldsymbol{{\cdot}}\left(\frac{{\rm d}^2}{{\rm d}r^2} + \frac{1}{r}\frac{{\rm d}}{{\rm d}r} - \frac{m^2}{r^2}\right)\tilde{\psi}_n^{(m)}(r;k) = 0, \end{gather}$$

(2.15b)$$\begin{gather}\boldsymbol{O}^{(k,m)}\boldsymbol{{\cdot}}\left(\frac{{\rm d}^2}{{\rm d}r^2} + \frac{1}{r}\frac{{\rm d}}{{\rm d}r} - \frac{m^2}{r^2}- k^2\right)\left(\frac{{\rm d}^2}{{\rm d}r^2} + \frac{1}{r}\frac{{\rm d}}{{\rm d}r} - \frac{m^2}{r^2}\right)\tilde{\xi}_n^{(m)}(r;k) = 0, \end{gather}$$

where the linear operator

(2.16)\begin{equation} \boldsymbol{O}^{(k,m)} \equiv \left[\lambda_m(k) + {\rm i}\,n\varOmega - \left(\frac{{\rm d}^2}{{\rm d}r^2} + \frac{1}{r}\frac{{\rm d}}{{\rm d}r} - \frac{m^2}{r^2} - k^2\right)\right]. \end{equation}

Equations (2.15a,b) are solved with the finiteness condition at $r\rightarrow 0$ in (2.6e) leading to

(2.17a,b)\begin{align} \tilde{\psi}_n^{(m)}(r;k) = \mathcal{A}_n {\mathrm{I}}_m(j_n r) + \mathcal{B}_n r^m,\quad\tilde{\xi}_n^{(m)}(r;k) = \mathcal{C}_n {\mathrm{I}}_m(j_n r) + \mathcal{D}_n {\mathrm{I}}_m(kr) + \mathcal{E}_n r^{m}, \end{align}

where $\mathcal {A}_n,\mathcal {B}_n,\mathcal {C}_n,\mathcal {D}_n$ and $\mathcal {E}_n$ are constants of integration, ${\mathrm {I}}_m({\cdot })$ is the $m$th-order modified Bessel function of first kind and $j_{n}^2 \equiv k^2 + ({\lambda _m(k) + \textrm {i}\,n\varOmega })/{\nu }$ with $Re\{j_n\} > 0$. The compatibility condition in (2.11) may be further simplified using (2.12a,b), the Floquet ansatz (2.14a,b) and the expressions in (2.17). The algebra for this is lengthy but eventually leads to a very simple relation, viz.

(2.18)\begin{equation} \mathcal{B}_n + k\mathcal{E}_n = 0\quad \forall\ n \ \in \mathbb{Z}. \end{equation}

The constants $\mathcal {B}_n$ and $\mathcal {E}_n$ appear only in the combination $\mathcal {B}_n + k\mathcal {E}_n$ in subsequent algebra and, thus, (2.18) may be used to eliminate these constants. Consequently, the only constants which survive in further analysis are $\mathcal {A}_n,\mathcal {C}_n,\mathcal {D}_n$ and $\mathcal {M}_n$ (see (2.14c). The Floquet ansatz in (2.14a,b) implies that the velocity components may be written as

(2.19)$$\begin{align} \left(u_r,u_{\theta},u_z\right) &= \displaystyle\sum_{n={-}\infty}^{\infty}\left(\tilde{u}_{r,n}(r)\cos(m\theta)\cos(kz),\tilde{u}_{\theta,n}(r)\sin(m\theta)\cos(kz),\tilde{u}_{z,n}(r)\cos(m\theta)\sin(kz)\right) \nonumber\\ &\quad \times\exp\left[\left({\rm i}\,n\varOmega + \lambda_m(k)\right)t\right], \end{align}$$

where the (complex) eigenmodes $\tilde {u}_{r,n}(r),\tilde {u}_{\theta,n}(r)$ and $\tilde {u}_{z,n}(r)$ are determined using expressions (2.17a,b) in (2.8a). These are

(2.20a–c)\begin{equation} \left. \begin{aligned} & \tilde{u}_{r,n}(r) = \frac{m}{r}{\mathrm{I}}_m(j_nr)\mathcal{A}_n + kj_n{\mathrm{I}}_m^{'}(j_nr)\mathcal{C}_n + k^2{\mathrm{I}}_m^{'}(kr)\mathcal{D}_n, \\ & \tilde{u}_{\theta,n}(r) ={-}\left\{j_n {\mathrm{I}}_m^{'}(j_n r)\mathcal{A}_n + \frac{km}{r}\left({\mathrm{I}}_m(j_nr)\mathcal{C}_n + {\mathrm{I}}_m(kr)\mathcal{D}_n\right)\right\}, \\ & \tilde{u}_{z,n}(r) ={-}\{ j_n^2{\mathrm{I}}_m(j_nr)\mathcal{C}_n + k^2{\mathrm{I}}_m(kr)\mathcal{D}_n\}, \end{aligned} \right\} \end{equation}

where prime indicates differentiation with respect to the argument, e.g. ${\mathrm {I}}_m^{'}(z)\equiv {\textrm {d}{\mathrm {I}}_m}/{\textrm {d}z}$ and so on. Note that despite the presence of terms of the form $1/r$ in expressions (2.20a,b), the velocity components do not diverge at the axis of the cylinder. This may be easily verified for the case $m > 0$ and the asymptotic form of ${\mathrm {I}}_m(z)$ for small $z$.

The boundary conditions in (2.6a–d) may now be simplified employing expressions (2.19) and (2.20a–c) to obtain linear algebraic equations in $\mathcal {A}_n$, $\mathcal {C}_n,\mathcal {D}_n$ and $\mathcal {M}_n$. The algebra is provided in supplementary material and we provide only the normal stress boundary condition below,

(2.21)\begin{align} &\left[\mu\left\{ k\mathcal{D}_n \left[ (k^2 - j_n^2)\dfrac{k{\mathrm{I}}_m'(kR_0)}{R_0} - \left(k^2 + j_n^2 + \dfrac{2m^2}{R_0^2}\right)k^2{\mathrm{I}}_m''(kR_0)\right] - 2 \left(k^2 + \frac{m^2}{R_0^2}\right)j_n^2{\mathrm{I}}_m''(j_nR_0)k\mathcal{C}_n\right.\right.\nonumber\\ &\quad \left.-\,2\left(k^2 + \dfrac{m^2}{R_0^2}\right)\dfrac{m}{R_0}\left(j_n{\mathrm{I}}_m'(j_nR_0) - \dfrac{{\mathrm{I}}_m(j_nR_0)}{R_0}\right)\mathcal{A}_n \right\} \nonumber\\ &\left.-\, \dfrac{T}{R_0^2}\left(k^2 + \frac{m^2}{R_0^2}\right)\left(k^2R_0^2 + m^2-1\right)\mathcal{M}_n\right]\left(\dfrac{2R_0^2}{\rho\left(k^2R_0^2 + m^2\right)}\right) = h\left[\mathcal{M}_{n - 1} + \mathcal{M}_{n + 1}\right]. \end{align}

Equation (2.21) is solved symbolically in Mathematica using expressions for $\mathcal {A}_n$, $\mathcal {C}_n$ and $\mathcal {D}_n$ in terms of $\mathcal {M}_n$ to obtain a single equation relating $\mathcal {M}_{n-1},\;\mathcal {M}_{n}$ and $\mathcal {M}_{n+1}$ for $n=0,1,2,3,\ldots, N$. Equation (2.21) is thus written as a generalized eigenvalue problem

(2.22)\begin{equation} \boldsymbol{A}\boldsymbol{{\cdot}}\boldsymbol{\mathcal{M}} = h \, \boldsymbol{Q}\boldsymbol{{\cdot}}\boldsymbol{\mathcal{M}} \quad {n = 0,1,2,\ldots,N}, \end{equation}

where $\boldsymbol {A}$ and $\boldsymbol {Q}$ are matrices and we have taken $N=30$ terms in the Fourier series for this study (see supplementary material). Expressing $\lambda _m(k) = \tilde {\mu } + {I}\alpha$, where ${I} \equiv \sqrt {-1}$, the sub-harmonic case is $\alpha =\varOmega /2$ and harmonic case is $(\alpha = 0)$ (Kumar & Tuckerman Reference Kumar and Tuckerman1994). With $\tilde {\mu }=0$, the resultant equations are solved using the Matlab generalized eigenvalue solver eig(,), MATLAB (2015) to obtain the stability boundaries on the wavenumber $k$ vs forcing $h$ plane for a given choice of $m$, forcing frequency $\varOmega$ and fluid parameters $T,\rho,\mu$ and $R_0$. The stability charts obtained from Floquet analysis will be discussed in § 4.

3.1. Inviscid limit of (3.6) and (3.7)

The first check on (3.7) is to demonstrate that it reduces to (2.3) (Matheiu equation on an inviscid cylinder) in the inviscid limit. In the inviscid limit, $l\rightarrow \infty$ (for fixed $s$) and it may be shown that $\lim _{\nu \rightarrow 0}\tilde {\zeta }(s)\rightarrow 0$ and $\lim _{\nu \rightarrow 0}\tilde {\chi }(s)\rightarrow 1$ in (3.7). For this, we have used the asymptotic expressions for ${\mathrm {I}}_{m}(z)$ and ${\mathrm {I}}_{m}^{'}(z)$ as $z\rightarrow \infty$ and fixed $m$ (Olver Reference Olver2021). Consequently, the inversion of (3.6) into the time domain becomes trivial leading to the Mathieu equation (Patankar et al. Reference Patankar, Farsoiya and Dasgupta2018) for potential flow,viz.

(3.8)\begin{equation} \frac{{\rm d}^2a_m}{{\rm d}t^2} + \frac{\mathrm{I}_m'(kR_0)}{\mathrm{I}_m(kR_0)}\left[ \dfrac{T}{\rho R_0^3}kR_0\left(k^2R_0^2 + m^2-1\right) + kh \cos\left(\varOmega t \right) \right] a_m(t) = 0, \end{equation}

where we have used $\mathcal {F}(r,t) = -h({r}/{R_0})\cos (\varOmega t)$ in writing (3.8).

3.2. Unforced ($h=0$) limit of (3.6)

The next test is to show that in the absence of forcing, expression (3.6) leads to the correct dispersion relation for free, viscous modes. We demonstrate this for the axisymmetric case where expressions for $\tilde {\chi }(s)$ and $\tilde {\zeta }(s)$ (see below (3.7)) are particularly very simple, viz. for $m=0$, we have

(3.9)\begin{equation} \tilde{\chi}(s) \rightarrow \frac{l^2-k^2}{l^2+k^2}=\frac{s}{s + 2\nu k^2},\quad \tilde{\zeta}(s) \rightarrow -\frac{k^2l}{l^2+k^2}\frac{{\mathrm{I}}_0^{''}(lR_0)}{{\mathrm{I}}_0^{'}(lR_0)} ={-}\frac{\nu lk^2}{s + 2\nu k^2}\frac{{\mathrm{I}}_0^{''}(lR_0)}{{\mathrm{I}}_0^{'}(lR_0)}. \end{equation}

These may be obtained from the observation that, for $m=0$, $\varLambda _1(s)$ diverges while $\varLambda _2(s)$ and $\varLambda _3$ remain finite. Using expressions 3.9 in (3.6) leads to

(3.10)$$\begin{gather} \left[s^2\tilde{a}_0 - sa(0)\right] + 2\nu k^2\frac{\mathrm{I}_0''(kR_0)}{\mathrm{I}_0(kR_0)}\left[ s\tilde{a}_0 - a(0)\right] -4\nu k \frac{\mathrm{I}_0'(kR_0)}{\mathrm{I}_0(kR_0)}\frac{\nu lk^2}{s + 2\nu k^2}\frac{{\mathrm{I}}_0^{''}(lR_0)}{{\mathrm{I}}_0^{'}(lR_0)}\left[ s\tilde{a}_0 - a(0)\right]\nonumber\\ + \frac{\mathrm{I}_0'(kR_0)}{\mathrm{I}_0(kR_0)}\frac{s}{(s + 2\nu k^2)}\left[ \frac{T}{\rho R_0^3} kR_0\left(k^2R_0^2-1\right) \tilde{a}_0\right] = 0, \end{gather}$$

implying that

(3.11)\begin{align} \tilde{a}_0(s;k) = \dfrac{\left[s +2\nu k^2\dfrac{{\mathrm{I}}_{0}^{''}(kR_0)}{{\mathrm{I}}_{0}(kR_0)} - \dfrac{4\nu^2 l k^3}{s + 2\nu k^2}\dfrac{{\mathrm{I}}_{0}^{'}(kR_0)}{{\mathrm{I}}_{0}(kR_0)}\dfrac{{\mathrm{I}}_0^{''}(lR_0)}{{\mathrm{I}}_0^{'}(lR_0)} \right]}{s^2 + 2\nu k^2\left\{\dfrac{{\mathrm{I}}_{0}^{''}(kR_0)}{{\mathrm{I}}_{0}(kR_0)} - \dfrac{2\nu l k}{s + 2\nu k^2}\dfrac{{\mathrm{I}}_{0}^{'}(kR_0)}{{\mathrm{I}}_{0}(kR_0)}\dfrac{{\mathrm{I}}_0^{''}(lR_0)}{{\mathrm{I}}_0^{'}(lR_0)}\right\}s - \dfrac{s}{s + 2\nu k^2} \sigma_0^2 }a(0). \end{align}

Comparing the denominator of (3.11) with expression (2.2b), and replacing $s\rightarrow \sigma$, we find that these are the same expressions. This is consistent as the viscous dispersion relation for free perturbations is obtained from the homogenous solution to the linear set of equations governing $\tilde {A}(s), \mathcal {\tilde {C}}(s),\mathcal {\tilde {D}}(s)$ and $\tilde {a}_m(s;k)$. The denominator of (3.11) represents the determinant of the homogenous part of these equations (Prosperetti Reference Prosperetti1976; Farsoiya, Roy & Dasgupta Reference Farsoiya, Roy and Dasgupta2020) and, thus, leads us to the dispersion relation provided in (2.2b). We have thus verified that (3.6) produces the correct dispersion relation in the unforced, axisymmetric limit.

3.3. Flat interface limit of (3.7)

We demonstrate that in the limit $R_0\rightarrow \infty$ (flat interface limit), our (3.6) reduces to the following equation ($\partial _t \equiv {\textrm {d}}/{\textrm {d}t}$) (Beyer & Friedrich Reference Beyer and Friedrich1995):

(3.12)$$\begin{gather} \left\{\frac{1}{k}{\left( \partial_t + 2\nu k^2 \right)}^2 + \left( \frac{Tk^2}{\rho} + h\cos\left(\varOmega t \right) \right)\right\}a_0(t) \nonumber\\ - \frac{4\nu^{3/2}k^2}{\rm \pi}\int_{-\infty}^{t} \sqrt{\frac{\rm \pi}{t - \tau}}\exp(-\nu k^2(t - \tau))\left(\partial_\tau + \nu k^2\right)a_0(\tau)\,{\rm d}\tau = 0. \end{gather}$$

The algebra for this is lengthy and is provided in Appendix B. Equation (3.12) is an analogue of (3.7) governing Faraday waves on a flat surface and was obtained by Beyer & Friedrich (Reference Beyer and Friedrich1995) (deep-water limit).

Having demonstrated the consistency of (3.6) and (3.7), we will return to analysing these at the end of § 5. Equation (3.7) is solved numerically in Mathematica using built-in numerical Laplace inversion subroutines (Wolfram Research, Inc. 2017) and results will be compared with DNS in § 5 in the context of RP stabilisation. In the next section we discuss the stability plots obtained from Floquet analysis which will suggest the RP stabilisation strategy.

4.1. Physical realisation of the radial, body force and comparison with experiments

The radial, oscillatory body force in (2.1) acts normal to the unperturbed surface of the cylinder mimicking a time-dependent gravitational force in the classical Faraday wave experiment (Benjamin & Ursell Reference Benjamin and Ursell1954). Such a harmonic body force of mechanical origin has been realised in the recent experiments of Maity et al. (Reference Maity, Kumar and Khastgir2020). These authors report observations of sub-harmonic, standing waves on liquid half-cylinders subject to substrate vibration. The experiments comprise a vertically vibrated, static, liquid half-cylinder (length $\approx 10$ cm, radius $\approx 0.3$ cm) with a pinned contact line. The vibration manifests as a body force on the cylinder in the oscillating frame of reference. Although this force is not purely radial, at high vibration frequency the azimuthal velocity produced due to the azimuthal force component is expected to be small. Consequently, the force may be treated as being purely radial and acting on a quiescent cylinder in a first approximation (see discussion in the first paragraph of page $7$, columns $1$ and $2$ in Maity et al. Reference Maity, Kumar and Khastgir2020).

Comparing their experimental results to inviscid linear stability predictions for the base state represented by (2.1), the authors report good agreement with the linear theory reported in (Patankar et al. Reference Patankar, Farsoiya and Dasgupta2018; Maity et al. Reference Maity, Kumar and Khastgir2020). Notably they (Maity et al. Reference Maity, Kumar and Khastgir2020) observe a sub-harmonic parametric instability also predicted by the theoretical model in (2.1). We will demonstrate that predictions from our present viscous Floquet analysis for the threshold acceleration of unstable, non-axisymmetric modes are in agreement with the experimental observations in Maity et al. (Reference Maity, Kumar and Khastgir2020). Despite this agreement, it should be noted that this is at best a qualitative comparison only. This is so because our (and Maity et al. Reference Maity, Kumar and Khastgir2020) theoretical model represented by (2.1) considers a full cylinder and does not explicitly account for the substrate present in the experiments. Interestingly, the choice of eigenmode in Patankar et al. (Reference Patankar, Farsoiya and Dasgupta2018) and the current study implies that if our full cylinder were to be treated as a half-cylinder on a hypothetical flat substrate at $\theta =(0,{\rm \pi} )$, the normal velocity by construction is zero at this substrate (see the expression for $u_{\theta }$ in (2.19)) and, thus, the non-penetration boundary condition at the substrate is automatically satisfied. However, the pinned contact line condition or the no-slip condition at the substrate cannot be respected. In contrast, the choice of the eigenmode in the inviscid study of Maity et al. (Reference Maity, Kumar and Khastgir2020) ensures that the pinned boundary condition at the (hypothetical) substrate is observed at $\theta = 0,{\rm \pi}$ ($\eta (z,\theta,t) \propto \sin (m\theta )$), but the no-penetration condition is then violated (their equation (4) implies $u_{\theta } \propto \cos (\theta )$). In the inviscid case, both choices of eigenmodes lead to an identical conclusion, viz. the amplitude of the perturbation is governed by the Mathieu equation. Note that equation (5) in Maity et al. (Reference Maity, Kumar and Khastgir2020) can be obtained from equation (3.7) in Patankar et al. (Reference Patankar, Farsoiya and Dasgupta2018) by setting gravity $g=0$ in the former and the ambient fluid density $\rho ^{{O}}=0$ in the latter. Consequently, we conclude that while the experimental template of Maity et al. (Reference Maity, Kumar and Khastgir2020) seems like a good candidate for realising an (approximately) radial oscillatory body force, appropriate modifications to it are necessary to minimise contact line and other substrate effects not present in the theory. These modifications are particularly important for realizing the theoretically predicted RP stabilisation of liquid cylinders sans any substrate that forms the focus of the current study. We discuss these modifications next and also compare our theoretical predictions with experimental results in Maity et al. (Reference Maity, Kumar and Khastgir2020).

In the absence of vertical vibration, a quiescent liquid half-cylinder with pinned contact line (Maity et al. Reference Maity, Kumar and Khastgir2020) is a static rivulet (Davis Reference Davis1980). For rivulets with vertical extent below the capillary length (as is the case for Maity et al. Reference Maity, Kumar and Khastgir2020), the effect of gravity is negligible and capillary effects dominate their shape and dynamics (Bostwick & Steen Reference Bostwick and Steen2018). It is useful to recall a few points concerning their stability as enumerated in Davis (Reference Davis1980) and, more recently, by Bostwick & Steen (Reference Bostwick and Steen2018). Importantly, the no-penetration and pinned boundary condition on the substrate are both explicitly accounted for in these studies. Linear analysis predicts that the classical RP instability of a liquid cylinder (that we seek to dynamically stabilise) persists for a static rivulet on a substrate, provided the contact angle exceeds ${\rm \pi} /2$ when the rivulet is unstable to pinned, varicose modes; see figure 5 in Davis (Reference Davis1980) and figure 4(a) in Bostwick & Steen (Reference Bostwick and Steen2018). As the experiments of Maity et al. (Reference Maity, Kumar and Khastgir2020) employ a half-cylinder with a contact angle of ${\rm \pi} /2$ (in the base state), it is expected that the RP instability mechanism does not operate for them where pinned varicose and sinuous modes are known to be stable (Bostwick & Steen Reference Bostwick and Steen2018). The instability that is expected in this case is parametric instability, when the amplitude of forcing exceeds a threshold. This expectation tacitly assumes that the stability predictions for the radially forced cylinder governed by (2.1) can (at least) qualitatively describe the parametric instability seen in these experiments (Maity et al. Reference Maity, Kumar and Khastgir2020). As will be seen in the following, this expectation receives support from experimental data.

As a first step it is verified that despite the length of the half-cylinder being more than its circumference in all cases (Maity et al. Reference Maity, Kumar and Khastgir2020), the authors do not report the RP instability in the absence of forcing, consistent with what is predicted theoretically (Bostwick & Steen Reference Bostwick and Steen2018). Instead, the authors detect parametric instability of non-axisymmetric modes ($m > 0$) at half the forcing frequency (Maity et al. Reference Maity, Kumar and Khastgir2020). In figure 4 we compare experimental data for the parametric instability in Maity et al. (Reference Maity, Kumar and Khastgir2020) with predictions from our viscous Floquet theory described earlier in § 2.3.1. We obtain acceleration thresholds $h$ relevant to non-axisymmetric modes (i.e. non-RP modes) on a radially forced, viscous liquid cylinder from our Floquet analysis. The modes (i.e. values of $(k,m)$) are taken from the experimental observations in Maity et al. (Reference Maity, Kumar and Khastgir2020) and it has been checked that the unstable modes manifested in their experiments are close to the minima of the first sub-harmonic tongue(s) at various forcing frequencies. Figure 4(a) shows a comparison of the threshold acceleration ($h_c$) at different frequencies from theory and experiments. A reasonably good agreement is apparent from the figure, despite significant differences between the theoretical model and the experimental set-up discussed earlier. These conclusions also concur with Maity et al. (Reference Maity, Kumar and Khastgir2020) who compared the wavenumbers $(k,m)$ of the most unstable mode at the inception of instability with their inviscid theory and reported good agreement.

Figure 4.

(a) Comparison of critical acceleration threshold obtained from Maity et al. (Reference Maity, Kumar and Khastgir2020) (their figure 7a,d) with present Floquet theory. We choose $\rho = 1, \mu = 0.01, R = 0.31, T = 72, g = 0$ in centimetre–gram–second (c.g.s.) units corresponding to water. In the data gap seen around $37$ Hz and $80$ Hz, the authors (Maity et al. Reference Maity, Kumar and Khastgir2020) do not report observing a single unstable wavelength but chaotic, irregular behaviour of the interface. The acceleration threshold on the vertical axis is expressed in units of gravity $g$ to aid intuition, although gravity is neglected in the model. The dimensions of the liquid half-cylinder in Maity et al. (Reference Maity, Kumar and Khastgir2020) are close to the air–water capillary length and, hence, gravity is expected to be negligible. (b) Suggested configuration for testing RP stabilisation by vertical oscillations.

Turning our attention to the RP mode stabilisation that is the focus of our study, it is evident that if the aforementioned experimental template (Maity et al. Reference Maity, Kumar and Khastgir2020) is to be used for testing the current RP stabilisation predictions (discussed in next section), the effect of the substrate needs to be minimised, if not eliminated. This can be achieved for a liquid cylinder on a substrate with contact angle approaching ${\rm \pi}$. The effect of increasing the contact angle far beyond ${\rm \pi} /2$ serves two purposes. On the one hand, it reduces liquid contact with the substrate, the configuration in the limit of the contact angle $\rightarrow {\rm \pi}$ resembling a cylinder on a wire support; see figure 4(b). Equally important, with increasing contact angle (above and beyond ${\rm \pi} /2$) the cylinder becomes susceptible to an increasingly wide range of RP unstable capillary modes and nearly approaches the free cylinder limit as the contact angle tends to ${\rm \pi}$ (the highest growth rates are observed at $\approx 135^{\circ }$, see figure 6 in Bostwick & Steen Reference Bostwick and Steen2018). Previous studies (Davis Reference Davis1980; Bostwick & Steen Reference Bostwick and Steen2018) have examined the RP unstable range for this configuration, predicting that the unstable wavenumber range gets somewhat diminished from $0 < kR_0 < 1$ in the free cylinder case to $0 < kR_0 < \sqrt {3/4}$ (Davis Reference Davis1980) for a cylinder on a wire support. We thus anticipate that the dynamic stabilisation proposed here for the idealised case of a free cylinder subject to a radial oscillatory body force could serve as a useful model to stabilise RP unstable modes of a liquid cylinder on a support with large contact angles in the range $135^{\circ }$–$180^{\circ }$ and minimal substrate contact. This is achieved by subjecting the latter to high-frequency vertical oscillations as depicted in figure 4(b). It should be borne in mind that a number of assumptions have been employed leading to the proposed set-up in figure 4(b). We assume that the oscillatory body force may be treated as being approximately radial, analogous to the half-cylinder case, even when the contact support is minimal. The inability to satisfy the no-slip condition at the substrate can be partially offset, by suitably coating it to induce slip (Haefner et al. Reference Haefner, Benzaquen, Bäumchen, Salez, Peters, McGraw, Jacobs, Raphaël and Dalnoki-Veress2015). However, other substrate boundary conditions cannot be satisfied, and due to the reduced substrate contact we assume that the effect of these are minimal. The validity of these assumptions and the resultant impact on the applicability of our present theoretical model to the proposed experiments are not intuitively obvious, and require systematic further investigation. We expect that the theoretical predictions made here will serve as a useful benchmark against which observations from the proposed experiments can be compared.

4.2. Linear stability predictions of RP mode stabilisation

Having demonstrated the feasibility of our theoretical model, we now shift attention to discussing the RP stabilisation strategy as suggested by the Floquet analysis described earlier. Referring to figure 5(a) (case 1 in table 3 provides the parameters), we wish to stabilise the axisymmetric RP unstable mode ($k_0=4.8,m_0=0)$ by subjecting the cylinder to an optimum forcing $h$. This cylinder is chosen to be of length $L=1.31$ cm or $2{\rm \pi} /k_0$. Thus, our stabilisation strategy ensures that the cylinder is as long as the longest RP unstable mode that we seek to stabilise. This ensures that shorter wavelength modes (both axisymmetric and three dimensional) are also simultaneously stable at the optimim level of forcing (see below). As shown in figure 5(a), the viscous stability tongues are moved upwards due to viscosity (Kumar & Tuckerman Reference Kumar and Tuckerman1994), compared with the inviscid tongues which touch the wavenumber axis (black dashed line in left panel). The figure shows that the critical threshold of forcing (we will call it $h_{{cr1}}$ hereafter) for stabilising ($k_0=4.8,m_0=0)$ is $h_{{cr1}}=1.23\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$, and the applied forcing ($h$) needs to satisfy $h > h_{{cr1}}$ for stabilisation of this mode. Simultaneously, we also need to ensure that $h$ is below a second threshold $h_{{cr2}}$. This second threshold ($h_{{cr2}}$) is chosen to be the ordinate corresponding to the lowest minima among all the stability tongues in figure 5(a,b). For stabilisation, we require $h_{{cr1}} < h_{{cr2}}$ and this is ensured by using the frequency of forcing $\varOmega$ as a control parameter for a given set of fluid parameters. Once we have chosen an $\varOmega$ which satisfies the ordering $h_{{cr1}} < h_{{cr2}}$, any choice of $h$ satisfying $h_{{cr1}} < h < h_{{cr2}}$ not only stabilises the primary mode ($k_0,m_0$) but also keeps moderately high modes ($k > k_0$ for $m=0,1,2,3,4\ldots$) stable.

Figure 5.

Stability plots for (a) axisymmetric ($m=0$) and (b) non-axisymmetric ($m=1,2,3,4$) modes with case 1 parameters, table 3 ($\varOmega = 600{\rm \pi}$). For $h > 0$, grey and white regions are unstable and stable, respectively. For panel (a), bold black lines $\rightarrow$ viscous tongue, black dashed line $\rightarrow$ inviscid tongue; inset shows the de-magnified view. The mode ($k_0=4.8,m_0=0$) is stabilised for $h > h_{{cr1}}=1.23\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$. The optimum forcing satisfies $h_{{cr1}} < h < h_{{cr2}}$ with $h_{{cr2}}=2.05\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$ for $m=4$ (see panel b). The chosen $h=1.8\times 10^4$ (indicated by a red symbol and solid red line in panels (a,b), respectively) keeps the cylinder stable. The $h_{{cr1}}$ scales as $\sim$ $({\varOmega }/{k})\sqrt {{T}/{\rho R_0^3}} \approx 2 \times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$ in comparison to capillary acceleration of an RP unstable mode, viz. $({T}/{\rho R_0^2}) \approx 500\, \textrm {cm}\,\textrm {s}^{-2}$, i.e. approximately $40$ times the latter. Results are shown for (a) $m=0$; (b) $m=1,2,3,4$.

Note that viscosity plays a very important role in this stabilisation as by displacing the (in)stability tongues upward, it allows for the possibility of choosing the forcing such that $h_{{cr1}} < h < h_{{cr2}}$. In the inviscid case, this is impossible to arrange as $h_{{cr}2}=0$ because in the inviscid case all (instability) tongues touch the wavenumber axis. Consequently, in an inviscid system if we force the cylinder at $h > h_{{cr}1}$, while the RP mode ($k_0,m_0=0$) is definitely stabilised, at long time (Patankar et al. Reference Patankar, Farsoiya and Dasgupta2018) higher modes (axisymmetric and non-axisymmetric) are produced due to nonlinearity and some of these are inevitably linearly unstable at the chosen level of forcing $h$. As a consequence, only a short-lived quasi-stabilisation is achieved in inviscid systems, thus rendering the stabilisation strategy unsuitable (this was shown in figure 3b). The situation is rectified by including viscosity into our analysis. Refer to figure 5 where the red dot in the left panel and the solid red line in the right panel indicates a suggested optimal value of $h$ satisfying $h_{{cr1}} < h < h_{{cr2}}$ for the RP mode $k_0=4.8,m_0=0$. Note that the high modes (i.e. those with $k \gg k_0$ and $m\gg m_0$) which can be generated due to nolinearity are also associated with high rates of dissipation. Consequently, we need not take into account the stability of very high modes in our stabilisation strategy. For the present purpose, we found it adequate to ensure that at the chosen value of $\varOmega$ and $h$, the primary mode $(k_0,m_0)$ as well as modes up to $(7k_0,\;m=0,1,2,3,4)$ are stable. This is found to be adequate for stabilisation of the liquid cylinder (at least up to the simulation time of several hundred forcing cycles).

An important point to note here is that although our theory has been developed assuming that a continuous range of RP modes with arbitrary long wavelengths ($k \rightarrow 0$) are accessible to our system, in practice there is a finite upper limit on the maximum wavelength that the system can access (due to axial confinement). In validating the present stability predictions via DNS (see § 5), we chose the length $L$ of the unperturbed cylinder to be $L={2{\rm \pi} }/{k_0}$, $k_0$ being the wavenumber of the axisymmetric RP unstable mode we intend to stabilise. Boundary conditions (periodic) in the axial ($z$) direction imply that only integral multiples of wavenumber $k_0$ are allowed to appear in our simulations. This ensures that wavenumbers verifying $k < k_0$ are not accessible to our system, although it is clear from figure 5(a) that such axisymmetric modes can continue to be unstable at the optimal level of forcing ($h=1.8\times 10^4$). We shall return to this point at the end of this study. For stabilising the mode ($k_0=4.8,m_0=0$), we have chosen $h=1.8\times 10^4$ (satisfying $h_{{cr}1} < h < h_{{cr}2}$) as indicated by the red dot in figure 5(a). It will be shown in § 5 through DNS that exciting the perturbation $k_0=4.8,m_0=0$ on the cylinder at $t=0$ with the forcing strength $h=1.23\times 10^4$ (at $\varOmega =600{\rm \pi}$) allows it to remain stable (at least up to the simulation time of several hundred forcing time periods). It will also be seen that the imposed perturbation decays to zero at long time, in excellent agreement with the solution to (3.7).

We next provide the optimal forcing strength for a slightly longer wavelength RP mode compared with the previous case. We choose to stabilise the axisymmetric RP unstable mode $(k_0=3.48,m_0=0)$. As described earlier, the cylinder length is now chosen to be $L = {2{\rm \pi} }/{k_0} = 1.81$ cm such that $k_0=3.48$ corresponds to the longest wave which can appear on this cylinder. This mode is indicated with a pink star in figure 5(a). It is seen that $h_{{cr}1}$ for this mode is $\approx 4.1\times 10^4\, \textrm {cm}\,\textrm {s}^{-2}$ and, thus, we do not satisfy $h_{{cr1}} < h_{{cr2}}$ (the minima of all the axisymmetric and non-axisymmetric stability tongues are much lower than $h_{{cr1}}$). Choosing simply $h > h_{{cr}1}$ allows the possibility of higher unstable modes to appear in simulations, as discussed in the last paragraph. In order to prevent this we now use the forcing frequency $\varOmega$ as a tuning parameter. In figure 6 we have increased $\varOmega =2200{\rm \pi}$ (from $600{\rm \pi}$ earlier), holding all fluid parameters at the same value as earlier (this is case $2$ in table 3). The advantage of doing so is visible in figure 6(a,b) where it is seen that by increasing $\varOmega$, we have the desired ordering. For our chosen mode $(k_0=3.48,m_0=0)$, we can see that $h_{{cr}1} \approx 1.52\times 10^5$ and $h_{{cr}2} \approx 1.74\times 10^5$ (obtained from the minima of the $m=4$ tongue shown in the right pane) and the desired ordering $h_{{cr1}} < h_{{cr2}}$ exists at this forcing frequency. The optimal level of forcing is chosen to be $h=1.65\times 10^5\, \textrm {cm}\,\textrm {s}^{-2}$ (indicated by the red dot and the solid red line in the left and right panel, respectively). It will be shown in the next section through DNS that this mode is also stabilised at this optimal forcing, at least up to the simulation time exceeding two thousand forcing time periods.

Figure 6.

Stability plots for (a) axisymmetric $m=0$ and (b) non-axisymmetric ($m=1,2,3,4$) modes with case 2 parameters, table 3 ($\varOmega =2200{\rm \pi}$). The figures are to be read in the same way as figure 5. The mode ($k_0=3.48,m_0=0$) is stabilised for $h > h_{{cr1}}=1.52\times 10^5\, \textrm {cm}\,\textrm {s}^{-2}$. The optimum forcing satisfies $h_{{cr1}} < h < h_{{cr2}}$. Here $h_{{cr2}}=1.74\times 10^5\, \textrm {cm}\,\textrm {s}^{-2}$ for $m=4$ (panel b). The chosen $h=1.65\times 10^5$ (indicated by a red symbol and solid red line in panel (a,b), respectively) keeps the cylinder stable. The $h_{{cr1}}$ scales as $\sim$ $({\varOmega }/{k})\sqrt {{T}/{\rho R_0^3}} \approx 1 \times 10^5\, \textrm {cm}\,\textrm {s}^{-2}$ in comparison to capillary acceleration of an RP unstable mode, viz. $({T}/{\rho R_0^2}) \approx 500\, \textrm {cm}\,\textrm {s}^{-2}$, i.e. approximately 200 times the latter. Results are shown for (a) $m=0$; (b) $m=1,2,3,4$.

5.1. Stabilisation of RP modes: DNS results and comparison with theory

Figure 8 shows stabilisation of the RP mode $k_0=4.8,m=0$ in DNS, both axisymmetric as well as three dimensional (refer to figure 5 for stability chart for this case). This is case $1$ in table 3 and shows stabilisation of the mode $k_0=4.8,m_0=0$ (subscripts $0$ are used for primary modes, viz. the modes excited initially in DNS). The dotted lines in red and blue are from DNS and nearly overlap. These indicate the amplitude of the interface as a function of time (the interface is tracked at the centre of the domain at $\theta =0$, see figure 7). The signals show stable, underdamped behaviour, decaying to zero after a few hundred forcing cycles ($\approx 400$ cycles). Note the excellent agreement between the DNS signals and the numerical solution to (3.7) indicated by the solid black line. We mention here that (3.7) in the axisymmetric limit reduces to (5.3) as discussed later in § 5. In all our comparison of DNS with analytical predictions that follow, we have solved (5.3) numerically in Mathematica (Wolfram Research, Inc. 2017) (this is referred to as analytical solution in all figure legends) and the script file for the same is provided as additional supplementary material. The inset to figure 8 shows that superposed on the long time underdamped oscillations are fine-scale oscillations arising from the high-frequency (compared with the growth rate of the RP mode) forcing imposed on the cylinder. Also shown in figure 8 are two more DNS signals, one with forcing $h > h_{{cr}1}$ and another with $h < h_{{cr}2}$. Both forcing levels are outside the optimum window $h_{{cr}1} < h < h_{{cr}2}$ and, thus, stabilisation is not achieved (see figure 5 for the optimum forcing window).

Figure 8.

Case $1$ in table 3: DNS time signal (red and blue dots) for ($k_0=4.8,m_0=0$) excited at $t=0$ and $h_{{cr1}} < h < h_{{cr}2}$; refer to stability plot in figure 5. Solution to (3.7) (black line); destabilisation seen in axisymmetric DNS when $h < h_{{cr}1}$ (pink line) and when $h > h_{{cr}2}$ (green line). Note the excellent agreement between solution to (3.7) and DNS at least up to the simulation time of $600$ forcing cycles ($\tilde {t} \equiv t\varOmega /2{\rm \pi}$). This is in contrast to inviscid simulations in figure 3(b) where for the same $k_0$, stabilisation is seen for only three forcing cycles.

Table 3.

Direct numerical simulation parameters (c.g.s. units).

In figure 9(a) we further validate the stabilisation obtained in figure 8 by turning off forcing at $\tilde {t}=485$ in DNS. It is seen that the interface destabilises in the absence of forcing indicating that forcing is crucial to the observed stabilisation.

Figure 9.

(a) Effect of turning-off forcing on RP mode stabilisation. This is the same mode as figure 8 with forcing turned off at $\tilde {t} = 485 \approx$ for DNS. Subsequently the RP unstable mode displays unbounded growth. (b) Case $2$ in table 3: DNS time signal for the mode ($k_0=3.48,m=0$). Stabilisation is seen at least up to the simulation time of $3000$ forcing cycles with excellent agreement between DNS (axisymmetric) and the solution to (3.7). Refer to stability chart in figure 6 for this case with frequency increased to $\varOmega =2200$ compared with case $1$. (a) Case 1 in table 3, (b) time signal.

In figure 9(b) we show stabilisation of the RP unstable mode $k_0=3.48,m_0=0$ (case $2$ in table 3). Recall from our discussion in the previous section that the frequency of forcing $\varOmega$ was increased to $2200{\rm \pi}$ for this case in order to satisfy the ordering $h_{{cr}1} < h < h_{{cr}2}$ (refer to figure 6 for stability chart for this case). The figure shows that stabilisation is achieved and sustained at least for a simulation time exceeding $3000$ forcing cycles when the perturbation decays to zero in an underdamped manner.

5.2. Damping and the memory term

We return in this section to a discussion of terms in (3.7) that appear due to viscosity, viz. the damping and the memory terms. These terms are physically easiest to interpret in the axisymmetric limit. It is shown in the supplementary material that in this limit, (3.7) reduces to

(5.3)\begin{align} \left. \begin{aligned} & \frac{{\rm d}^2a_0}{{\rm d}t^2} + 2\nu k^2\left(1+ \frac{\mathrm{I}_0''(kR_0)}{\mathrm{I}_0(kR_0)}\right)\frac{{\rm d}a_0}{{\rm d}t} + \frac{\mathrm{I}_0'(kR_0)}{\mathrm{I}_0(kR_0)}\left[ \frac{T}{\rho R_0^3}kR_0 (k^2R_0^2-1)+ kh \cos(\varOmega t) \right] a_0(t) \\ & \qquad +\frac{4\nu^2 k^4}{\mathrm{I}_0(kR_0)}\int_{0}^{t}\hat{\boldsymbol{L}}^{{-}1}\left[\mathcal{K}(s)\right]\frac{{\rm d}a_0}{{\rm d}\tau}(t - \tau) \,{\rm d}\tau = 0,\\ & \text{where} \quad \mathcal{K}(s) = \left( \frac{\mathrm{I}_0''(kR_0)}{s} -\frac{l}{k}\frac{\mathrm{I}_0'(kR_0){\mathrm{I}}_0^{''}(lR_0)}{s{\mathrm{I}}_0^{'}(lR_0)}\right). \end{aligned} \right\} \end{align}

If we temporarily disregard the memory term in (5.3), then it is clear that the rest of the equation constitutes a damped Mathieu equation, i.e. the damped version of (2.3) for $m=0$ (axisymmetric). This is

(5.4)\begin{align} \frac{{\rm d}^2a_0}{{\rm d}t^2} + 2\nu k^2\left(1+ \frac{\mathrm{I}_0''(kR_0)}{\mathrm{I}_0(kR_0)}\right)\frac{{\rm d}a_0}{{\rm d}t} + \frac{\mathrm{I}_0'(kR_0)}{\mathrm{I}_0(kR_0)}\left[ \frac{T}{\rho R_0^3}kR_0 (k^2R_0^2-1)+ kh \cos(\varOmega t) \right] a_0(t) = 0. \end{align}

Equation (5.4) is the cylindrical analogue of its Cartesian counterpart which has been discussed in Kumar & Tuckerman (Reference Kumar and Tuckerman1994); Cerda & Tirapegui (Reference Cerda and Tirapegui1998) for viscous Faraday waves over a flat interface (see equation (4.21) in Kumar & Tuckerman (Reference Kumar and Tuckerman1994) or equation (3.4) in Cerda & Tirapegui Reference Cerda and Tirapegui1998). In order to put this analogy on a sound footing, we take the limit $R_0\rightarrow \infty$ (for fixed $k$) on (5.4) expecting to recover results relevant to a flat interface (as $R_0\rightarrow \infty$, the cylinder locally becomes flat). Using the identity $\lim _{z\rightarrow \infty }{\mathrm {I}}_{0}^{''}(z)/{\mathrm {I}}_{0}(z) = 1$, it is seen that the coefficient of the second term in 5.4 in this limit reduces to the damping coefficient of viscous capillary waves (deep water) on a flat interface, viz. $4\nu k^2$, which is the same as estimated in Kumar & Tuckerman (Reference Kumar and Tuckerman1994); Cerda & Tirapegui (Reference Cerda and Tirapegui1998). Note that the damping factor $4\nu k^2$ for a flat interface is obtained by estimating dissipation for potential flow (Kumar & Tuckerman Reference Kumar and Tuckerman1994). By analogy it may similarly be expected that the pre-factor $2\nu k^2(1+ {\mathrm {I}_0''(kR_0)}/{\mathrm {I}_0(kR_0)})$ in (5.4) arises from the damping of potential flow (Patankar et al. Reference Patankar, Farsoiya and Dasgupta2018) in the liquid cylinder. It has been verified that this is correct and the factor $2\nu k^2(1+ {\mathrm {I}_0''(kR_0)}/{\mathrm {I}_0(kR_0)})$ indeed agrees with the damping predicted by the coefficient of $\sigma$ in the dispersion relation in equation (5.10) of Wang, Joseph & Funada (Reference Wang, Joseph and Funada2005), which was obtained through a viscous potential flow calculation (VCVPF in their terminology with a crucial viscous pressure correction)

Turning now to the memory term in (5.3), we note that it does not depend on the forcing strength $h$. Thus, it persists even in the unforced limit ($h\rightarrow 0$), in which case (5.3) becomes one governing free perturbations. This equation was derived earlier by Berger (Reference Berger1988) by solving the corresponding IVP with $h=0$ and we have verified that the unforced limit of (5.3) agrees with the equation of Berger (Reference Berger1988) (see supplementary material). The Laplace inversion of $\mathcal {K}(s)$ in (5.3) is analytically feasible and may be expressed as an infinite summation over integrals from residue theory (see expression 79 in Berger Reference Berger1988). For convenience, we reproduce this here as the term on the right-hand side of (5.5) (the damping term in (5.5) has been slightly modified from Berger (Reference Berger1988) but is exactly equivalent to his expression),

(5.5)\begin{align} & \frac{{\rm d}^2a_0}{{\rm d}t^2} + 2\nu k^2\left(1+ \frac{\mathrm{I}_0''(kR_0)}{\mathrm{I}_0(kR_0)}\right)\frac{{\rm d}a_0}{{\rm d}t}(t) + \left[\frac{T}{\rho R_0^3} kR_0 \left(k^2R_0^2-1\right)\frac{{\mathrm{I}}_1(kR_0)}{{\mathrm{I}}_0(kR_0)}\right]a_0(t) \nonumber\\ &\quad = \frac{8\nu^2k^3}{R_0}\frac{{\mathrm{I}}_0(kR_0)}{{\mathrm{I}}_1(kR_0)}\int_{0}^{t}\,{\rm d}t^{'}\;\frac{{\rm d}a_0(t')}{{\rm d}t^{'}}\exp\left(-\nu k^2(t-t^{'})\right)\displaystyle\sum_{j_n}\frac{\exp\left[-\left(\dfrac{\nu}{R_0^2}\right)j_n^2(t-t^{'})\right]}{1 + \left(\dfrac{R_0k}{j_n}\right)^2}, \end{align}

where $j_n$ represents the $n$th (non-zero) zero of $J_1(j_n)=0$ (Berger Reference Berger1988). The origin of the infinite summation in (5.5) may be rationalised as follows: the initial condition of zero vorticity and surface deformation (i.e. $\eta (z,\theta,0)=a_0\cos (k_0z)$) excites all modes in the spectrum, (viz. two capillary modes and a countable infinite set of hydrodynamic modes García & González Reference García and González2008). The excitation of the countably infinite set of hydrodynamic modes (which are all purely damped modes) produces the infinite summation in the analytical expression for $a_0(t;k)$, also manifesting as the memory term(s) in (5.5). These conclusions for free perturbations on a cylinder have analogues on a flat surface (e.g. see equation (2.30) in Cerda & Tirapegui (Reference Cerda and Tirapegui1998) which expresses the amplitude as a sum over two capillary modes and an infinite sum over the hydrodynamic modes).

Physically, the presence of the memory term implies that the damping seen in DNS contains contributions not only from the potential part of the flow (as is modelled correctly by the damped Mathieu equation (5.4)) but also from the memory term(s) which arise due to the boundary layer at the free surface. We find that the contribution of the memory term in (5.4) increases as the kinematic viscosity of the fluid is increased, and is the largest (in the axisymmetric limit being studied here) when viscosity is sufficiently large for the stabilised response of the liquid cylinder to be overdamped. Figure 10(b) depicts this for the RP mode $k_0=4.8,m_0=0$ (case $3$ in table 3), highlighting the difference between the solution to the damped Mathieu equation (5.4) and the integro-differential equation (5.3). It is seen that at intermediate time ($80 < \tilde {t} < 100$), the damped Mathieu equation (5.4) underpredicts the damping that is seen in the DNS and in (5.3). The corresponding stability chart with the optimal level of forcing for stabilisation is indicated in the upper panel of figure 10(a).

Figure 10.

(a) Stability diagram for case $3$ in table 3. The viscosity has been doubled for this case compared with case $1$ in table 3. The RP mode $k_0=4.8,m_0=0$ and moderately higher modes are stabilised if $h_{{cr}1} < h < h_{{cr2}}$. Here $h_{{cr1}}=1.24\times 10^4$, and $h_{{cr2}}=3.74\times 10^4$ is determined from the non-axisymmetric stability plot for $m=4$ (not shown here). We choose $h=1.8\times 10^4$ for stabilisation as indicated by the red dot. (b) Time signal from axisymmetric DNS showing stabilisation for the RP unstable mode indicated by a red dot in (a), viz. $k=4.8,m=0$. Note the overdamped response and the excellent agreement with the solution to (3.7). Blue line: solution to the damped Mathieu equation (5.4). The analytical response is the solution to (5.3). (a) Result shown for $m=0$. (b) Time signal.

We conclude this study with a discussion on the limitation of the present stabilisation technique, viz. that it does not stabilise the entire RP unstable spectrum at any finite level of forcing, but it does stabilise all modes accessible to the cylinder (both axisymmetric and three dimensional) with $k > k_0$ on a cylinder of length $L = {2{\rm \pi} }/{k_0}$. This arises from the infinitely long cylinder assumption that we have made allowing all modes from $0 < k_0 < \infty$ to be present. In practice we expect to encounter liquid cylinders of finite length typically confined between supports. The boundary conditions at the end points (e.g. pinned, see Sanz Reference Sanz1985) can substantially modify the nature of the eigenmodes in the $z$ direction compared with the Fourier modes that we have assumed here.